Three-Dimensional Superconformal Field Theory, Chern-Simons Theorv, and Their Correspondence

Citation

Chung, Hee Joong (2014) Three-Dimensional Superconformal Field Theory, Chern-Simons Theorv, and Their Correspondence. Dissertation (Ph.D.), California Institute of Technology. doi:10.7907/7CA7-9C79. https://resolver.caltech.edu/CaltechTHESIS:06062014-152022421

Abstract

In this thesis, we discuss 3d-3d correspondence between Chern-Simons theory and three-dimensional N = 2 superconformal field theory. In the 3d-3d correspondence proposed by Dimofte-Gaiotto-Gukov information of abelian flat connection in Chern-Simons theory was not captured. However, considering M-theory configuration giving the 3d-3d correspondence and also other several developments, the abelian flat connection should be taken into account in 3d-3d correspondence. With help of the homological knot invariants, we construct 3d N = 2 theories on knot complement in 3-sphere for several simple knots. Previous theories obtained by Dimofte-Gaiotto-Gukov can be obtained by Higgsing of the full theories. We also discuss the importance of all flat connections in the 3d-3d correspondence by considering boundary conditions in 3d N = 2 theories and 3-manifold.

Item Type:	Thesis (Dissertation (Ph.D.))
Subject Keywords:	Three-Dimensional Superconformal Field Theory, Chern-Simons Theory, 3d-3d Correspondence
Degree Grantor:	California Institute of Technology
Division:	Physics, Mathematics and Astronomy
Major Option:	Physics
Thesis Availability:	Public (worldwide access)
Research Advisor(s):	Gukov, Sergei (advisor) Schwarz, John H. (co-advisor)
Group:	Caltech Theory
Thesis Committee:	Gukov, Sergei (chair) Schwarz, John H. (co-chair) Kapustin, Anton N. Porter, Frank C.
Defense Date:	30 May 2014
Record Number:	CaltechTHESIS:06062014-152022421
Persistent URL:	https://resolver.caltech.edu/CaltechTHESIS:06062014-152022421
DOI:	10.7907/7CA7-9C79
Default Usage Policy:	No commercial reproduction, distribution, display or performance rights in this work are provided.
ID Code:	8496
Collection:	CaltechTHESIS
Deposited By:	Hee Joong Chung
Deposited On:	10 Jun 2014 18:51
Last Modified:	01 Jul 2025 16:31

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Three-Dimensional Superconformal Field Theory, Chern-Simons Theory, and Their Correspondence Thesis by Hee Joong Chung In Partial Fulfillment of the Requirements for the Degree of Doctor of Philosophy California Institute of Technology Pasadena, California 2014 (Defended May 30, 2014) ii c 2014 Hee Joong Chung All Rights Reserved iii Abstract In this thesis, we discuss 3d-3d correspondence between Chern-Simons theory and three-dimensional N = 2 superconformal field theory. In the 3d-3d correspondence proposed by Dimofte-GaiottoGukov information of abelian flat connection in Chern-Simons theory was not captured. However, considering M-theory configuration giving the 3d-3d correspondence and also other several developments, the abelian flat connection should be taken into account in 3d-3d correspondence. With help of the homological knot invariants, we construct 3d N = 2 theories on knot complement in 3-sphere S 3 for several simple knots. Previous theories obtained by Dimofte-Gaiotto-Gukov can be obtained by Higgsing of the full theories. We also discuss the importance of all flat connections in the 3d-3d correspondence by considering boundary conditions in 3d N = 2 theories and 3-manifold. iv Acknowledgments I truly appreciate my advisors Sergei Gukov and John H. Schwarz for their support, guide, advice, patience, and encouragement. I learned physics and mathematics a lot from discussion with them. I also would like to thank Anton Kapustin for being my committee member and his inspiring lectures on quantum field theory and advanced mathematical methods in physics which I took in early stage of my graduate study. I am also grateful to Frank Porter for being my committee member for thesis defense and the candidacy exam, and it was my pleasure to work for him as a teaching assistant several times. I also appreciate Hirosi Ooguri for being a committee member for the candidacy exam and the reading course which I took in my first year. I am also very grateful to Tudor Dimofte and Piotr Sulkowski for collaboration. I have benefited from discussion with them. I would like to thank to fellow graduate students and postdocs at Caltech theory group including Miguel Bandres, Denis Bashkirov, Kimberly Boddy, Tudor Dimofte, Ross Elliot, Abhijit Gadde, Lotte Hollands, Yu-tin Huang, Matt Johnson, Christoph Keller, Hyungrok Kim, Arthur Lipstein, Kazunobu Maruyoshi, Joseph Marsano, Yu Nakayama, Tadashi Okazaki, Chan Youn Park, ChangSoon Park, Du Pei, Pavel Putrov, Ingmar Saberi, Sakura Schafer-Nameki, Grant Remmen, Kevin Setter, Jaewon Song, Bogdan Stoica, Kung-Yi Su, Meng-Chwan Tan, Heywood Tam, Ketan Vyas, Brian Willett, Itamar Yaakov, Masahito Yamazaki, Wenbin Yan, and Jie Yang. I am also grateful to Carol Silberstein for her help and assistance. It is also my pleasure to thank people outside Caltech, including Dongmin Gang, Seungjoon Hyun, Hee-Cheol Kim, Youngjoon Kwon, Kimyeong Lee, Sukyoung Lee, Sungjay Lee, Satoshi Nawata, Daniel Park, Jaemo Park, Piljin Yi, and Peng Zhao. Finally, I deeply appreciate my father, mother, and brother for their constant support and encouragement. v Contents Abstract iii Acknowledgments iv 1 Introduction 1 2 3d N = 2 Superconformal Theories 3 2.1 Aspects of 3d N = 2 supersymmetric gauge theories . . . . . . . . . . . . . . . . . . 3 3d N = 2 supersymmetry algebra and its irreducible representations . . . . . 4 2.1.1.1 3d N = 2 supersymmetry algebra . . . . . . . . . . . . . . . . . . . 4 2.1.1.2 Irreducible representations . . . . . . . . . . . . . . . . . . . . . . . 4 2.1.2 Supersymmetric actions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5 2.1.3 Parity anomaly . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6 2.1.4 Monopole operators and vortices . . . . . . . . . . . . . . . . . . . . . . . . . 6 2.1.5 3d N = 2 mirror symmetry . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7 2.1.1 2.2 Compactification of 3d N = 2 gauge theories on S 1 . . . . . . . . . . . . . . . . . . . 8 2.3 The exact calculations in 3d N = 2 superconformal theories . . . . . . . . . . . . . . 9 2.3.1 Supersymmetric partition function on the squashed 3-sphere . . . . . . . . . 10 2.3.2 Superconformal index . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 11 Holomorphic block in 3d N = 2 theories . . . . . . . . . . . . . . . . . . . . . . . . . 12 2.4.1 Holomorphic block as the BPS index . . . . . . . . . . . . . . . . . . . . . . . 13 2.4.2 Fusion of holomorphic and anti-holomorphic block . . . . . . . . . . . . . . . 13 2.4.3 Holomorphic block from supersymmetric quantum mechanics . . . . . . . . . 14 Sp(2L, Z) action on 3d conformal field theories . . . . . . . . . . . . . . . . . . . . . 15 2.4 2.5 3 Chern-Simons Theory 3.1 18 Analytic continuation of Chern-Simons theory . . . . . . . . . . . . . . . . . . . . . . 19 3.1.1 Structure of Chern-Simons partition functions . . . . . . . . . . . . . . . . . . 20 3.1.2 Wilson loop operator . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 22 vi 3.2 3.3 b Classical moduli space, A-polynomial and quantum A-polynomial of knot . . . . . . 22 3.2.1 Classical moduli space and A-polynomial . . . . . . . . . . . . . . . . . . . . 22 3.2.2 b Quantization and quantum A-polynomial . . . . . . . . . . . . . . . . . . . . 24 3.2.3 Recursion relation for Jones polynomial . . . . . . . . . . . . . . . . . . . . . 25 Ideal hyperbolic tetrahedra and their gluing . . . . . . . . . . . . . . . . . . . . . . . 26 3.3.1 An ideal tetrahedron, boundary phase space, and Lagrangian submanifold . . 26 3.3.2 Gluing tetrahedra and boundaries of 3-manifold . . . . . . . . . . . . . . . . . 28 3.3.3 2-3 Pachner move . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 30 3.3.4 Quantization, wavefunction, and partition function . . . . . . . . . . . . . . . 30 4 3d-3d Correspondence from Gluing Tetrahedra 34 4.1 Motivation for 3d-3d correspondence . . . . . . . . . . . . . . . . . . . . . . . . . . . 35 4.2 3d-3d correspondence of Dimofte-Gaiotto-Gukov . . . . . . . . . . . . . . . . . . . . 36 4.2.1 3d-3d correspondence for a tetrahedron . . . . . . . . . . . . . . . . . . . . . 36 4.2.2 Gluing procedures . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 38 Examples . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 39 4.3 4.3.1 4.3.2 SQED with Nf = 1, XYZ model, and 3d N = 2 mirror symmetry . . . . . . 39 4.3.1.1 SQED Nf = 1 and XYZ model from gluing tetrahedra . . . . . . . . 39 4.3.1.2 Holomorphic block, partition function on Sb3 , and index . . . . . . . 41 4.3.1.3 Supersymmetric parameter space . . . . . . . . . . . . . . . . . . . . 43 Trefoil knot and figure-eight knot . . . . . . . . . . . . . . . . . . . . . . . . . 44 5 Toward a Complete 3d-3d Correspondence 5.1 3d-3d correspondence revisited . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 46 5.1.1 M-theory perspective . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 46 5.1.1.1 Symmetries of T [M3 ] from the perspective of M5-brane system . . . 47 5.1.1.2 Supersymmetric parameter space . . . . . . . . . . . . . . . . . . . . 48 Brief comments on recent developments in 3d-3d correspondence . . . . . . . 48 5.1.2 5.2 46 Goal and strategy . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 49 5.2.1 T [M3 ] from homological knot invariants and Higssing . . . . . . . . . . . . . 49 5.2.2 Boundary of M3 and T [M3 ] . . . . . . . . . . . . . . . . . . . . . . . . . . . . 50 5.3 Contour integrals for Poincaré polynomials . . . . . . . . . . . . . . . . . . . . . . . 51 5.4 Knot polynomials as partition functions of T [M3 ] . . . . . . . . . . . . . . . . . . . . 55 5.4.1 Recursion relations for Poincaré polynomials . . . . . . . . . . . . . . . . . . 56 5.4.2 3d N = 2 gauge theories for unknot, trefoil knot, and figure-eight knot . . . . 58 Vortices in S ×q S . . . . . . . . . . . . . . . . . . . . . . . . . . . 63 The t = −1 limit and DGG theories . . . . . . . . . . . . . . . . . . . . . . . . . . . 69 5.4.3 5.5 2 1 and Sb3 . vii 5.6 5.5.1 The DGG theories . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 70 5.5.2 Indices and residues . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 72 5.5.3 Critical points and missing vacua . . . . . . . . . . . . . . . . . . . . . . . . . 73 5.5.4 Relation to colored differentials . . . . . . . . . . . . . . . . . . . . . . . . . . 75 Boundaries in three dimensions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 77 5.6.1 77 5.6.2 Bibliography Cutting and gluing along boundaries of M3 . . . . . . . . . . . . . . . . . . . 1 5.6.1.1 Compactification on S and branes on the Hitchin moduli space . . 79 5.6.1.2 Lens space theories and matrix models . . . . . . . . . . . . . . . . 80 5.6.1.3 Dehn surgery . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 84 Boundary conditions in 3d N = 2 theories . . . . . . . . . . . . . . . . . . . . 86 88 1 Chapter 1 Introduction In M-theory, there are M2-brane and M5-brane. The worldvolume theory of multiplet M2-branes is relatively well-understood by Bagger-Lambert and Aharony-Bergman-Jafferis-Maldacena. However, the worldvolume theory of multiple M5-brane is still very mysterious. From the string/M-theory argument, it is known that there exist 6d (2, 0) superconformal field theory, though people don’t know a proper description of this theory yet. Instead, physicsists have tried to compactify 6d (2, 0) theory on manifolds whose dimension is less than 6. For example, if one compactifies 6d (2, 0) theory on a circle, one obtains 5d maximally supersymmetric field theory. Or if one compactify 6d (2, 0) theory on the torus T 2 , 4d N = 4 superconformal theory is obtained. This program of compactification on general punctured Riemann surfaces C was pioneered by Gaiotto. He constructed 4d N = 2 superconformal field theory labelled by Riemann surface C. Information of Riemann surface is realized in 4d gauge theory side, for example, mapping class group of C is identified as S-duality group of 4d N = 2. More surprisingly, it turned out that instanton partition function pioneered by Nekrasov is actually identified with the conformal block of Liouville theory (or more generally Toda theory) on Riemann surface C. This surprising relation was found by Alday-Gaiotto-Tachikawa (AGT) and often called AGT correspondence or 2d-4d correspondence. Therefore from 2d-4d correspondence, one might expect that compactification of 6d (2, 0) theory on a certain d-dimensional manifold Md would give a certain 6 − d-dimensional superconformal field theory. About two years later, Dimofte-Gaiotto-Gukov (DGG) showed that it also holds for d = 3 by using gluing ideal hyperbolic tetrahedra. The 3d-3d correspondence states that when M5 branes are wrapped on a 3-manifold − M3 − we have 3d N = 2 superconformal field theories (SCFT) − T [M3 ] − described as the IR fixed points of abelian Chern-Simons-matter theories determined by M3 and other extra information. At the same time, physical quantities such as partition functions of non-supersymmetric complex Chern-Simons (CS) theory on M3 are matched with those of 3d N = 2 abelian superconformal theory. 2 Though it seemed successful, actually it was not a complete correspondence as also mentioned in the original papers. In the correspondence of DGG, abelian branch of flat connections in ChernSimons theory is lost. However, considering M-theory point of view and several developments, the 3d-3d correspondence should capture all branches of flat connections. In author’s work with T. Dimofte, S. Gukov, P. Sulkowski [1], we found 3d N = 2 theories cor- responding to knot complement in S 3 for several simple knots whose partition functions are equal to the homological knot invariants. Our examples give a full complete 3d-3d correspondence in the sense that they capture all flat connection without anything lost. It is also possible to reproduce theories constructed by DGG by Higgs mechanism from our theories. In addition, we see the importance of abelian flat connections by considering boundaries of 3d N = 2 theories and 3-manifold. This thesis has following structure; In Chapter 2, we review the basic aspects of 3d N = 2 supersymmetric theories with focus on the necessary elements in 3d-3d correspondence. In Chapter 3, we review the analytic continuation of Chern-Simons theory with focus on the relevant materials for our purpose. In Chapter 4, we summarize the 3d-3d correspondence by Dimofte-Gaiotto-Gukov obtained by gluing ideal hyperbolic tetrahedra. In Chapter 5, we engineer the 3d N = 2 theories corresponding to a knot complement in S 3 for unknot, trefoil knot, and figure-eight knot. We discuss Higgs mechanism to reproduce knot polynomials and the previous theories of DGG. We also see the importance of including all flat connection in 3d-3d correspondence by considering boundaries of 3d N = 2 theories and those of 3-manifold. 3 Chapter 2 3d N = 2 Superconformal Theories In this chapter, we review topics in 3d N = 2 theories with focus on the 3d-3d correspondence. In section 2.1, we summarize several aspects of 3d N = 2 theories including supersymmetry algebra, supersymmetric action, parity anomaly, monopole operators, vortices, and 3d N = 2 mirror symmetry. This section is mostly for setting up basic terminology or background for the rest of thesis. In section 2.2, we review aspects of compacitification of 3d N = 2 on S 1 . The resulting theory becomes effective 2d N = (2, 2) theory which is characterized by twisted superpotential. In section 2.3, we summarize results on the exact calculation for partition functions on the squashed 3-sphere Sb3 and superconformal index on S 2 ×q S 1 of 3d N = 2 theories. In section 2.4, we review aspects of holomorphic block in 3d N = 2 theories with discrete, massive, sueprsymmetric vacua. A partition function on the squashed 3-sphere Sb3 and superconformal index S 2 ×q S 1 can be understood as appropriate fusion of holomorphic and anti-holomorphic block. We mainly discuss formal aspects, so discussion is a bit abstract. Examples are deferred to section 4. In section 2.5, we summarize the Sp(2L, Z) action on 3d (super)conformal field theories. This will be eventually related to a certain similar symplectic action in 3-manifold side. 2.1 Aspects of 3d N = 2 supersymmetric gauge theories In this section, we review some basic aspects of 3d N = 2 supersymmetric gauge theories [2, 3]. 4 2.1.1 3d N = 2 supersymmetry algebra and its irreducible representations 2.1.1.1 3d N = 2 supersymmetry algebra 3d N = 2 supersymmetry algebra is obtained from the dimensional reduction of 4d N = 1 supersymmetry algebra, and it is given by {Qα , Qβ } = {Qα , Qβ } = 0, µ {Qα , Qβ } = 2σαβ Pµ + 2iαβ Z. (2.1) Here, α and β are indices for spinor which go from 1 to 2, and Z is a central charge which can be thought as the fourth component of momentum P in 4d. There is a U (1)R symmetry which rotates Q and Q in opposite phase, more explicitly we choose convention that Q is charged −1 and Q is charged +1 under U (1)R . 2.1.1.2 Irreducible representations The irreducible representation of 3d N = 2 superalgebra contains vector multiplet, chiral multiplet, which can also be obtained from dimensional reduction of 4d N = 1 multiplets. 3d N = 2 vector multiplet V consists of a gauge field Aµ , a complex Dirac spinor λ, a real scalar field σ, and a real auxiliary scalar D. In superspace with Wess-Zumino gauge, vector multiplet V is given by 1 V = −θσ µ θAµ (x) − θθσ + iθθθλ(x) − iθθθλ(x) + θθθθD(x) 2 (2.2) where σ can be thought as the fourth component of 4d gauge field. As in 4d, we also have chiral/anti-chiral field strength multiplets, 1 Wα = − DDe−V Dα eV , 4 1 W α = − DDe−V Dα eV 4 (2.3) The lowest component of a field strength mutliplet is a gaugino. It is also possible to define a multiplet whose lowest component is scalar field. This is called as linear multiplet, which is defined as α Σ = D Dα V, (2.4) satisfying α Dα Dα Σ = D Dα Σ = 0, Σ = Σ† (2.5) 3d N = 2 chiral multiplet Φ consists of a complex Dirac spinor ψ, a real scalar field φ, and a 5 real auxiliary scalar field F ; √ Φ=φ+ 2θψ + θθD, (2.6) and similarly for anti-chiral multiplet. 2.1.2 Supersymmetric actions The classical Yang-Mills kinetic terms for vector multiplet can be written in terms of superfield; 1 Lkin = 2 g Z d2 θTrWα2 + h.c. (2.7) where trace is over fundamental representation in this subsection. This can be also written in terms of linear multiplet; 1 Lkin = 2 g Z 1 d2 θd2 θ Tr Σ2 . 4 (2.8) The classical kinetic terms for chiral multiplets Φ in representation of gauge group is written as XZ Lkin,matter = d2 θd2 θ Φ†i eV Φi (2.9) i We can get mass terms in Lagrangian from non-zero vacuum expectation value (vev) of scalar component of background vector multiplet. If a background vector multiplet contains mθθ from a nonzero vev of an adjoint scalar and other components are turned off, we have Lmass,matter = XZ d2 θd2 θ Φ†i emθθ Φi , (2.10) i which gives a mass to matter multiplets. This mass is called as “real mass”. We also have Chern-Simons therm, which is given by LCS = k 4π Z d2 θd2 θ Tr Σ V (2.11) where k ∈ Z is a Chern-Simons level. For U (1) factor of gauge group, we can also add Fayet-Iliopoulos (FI) term; LF I = Z d2 θd2 θ ζ V (2.12) 6 where ζ is an FI parameter. This term can also be thought as LF I = Z d2 θd2 θ Σ V (2.13) where a linear multiplet Σ has a scalar component σ = ζ and the rest components are turned off. For abelian vector multiplets, more generally, CS term and FI term can be thought as a special case of LBF = kij 4π Z d2 θd2 θ Σi Vj . (2.14) For i 6= j, it is called as BF term in literature. In this thesis, we call it as mixed Chern-Simons term. 2.1.3 Parity anomaly Unlike 4d gauge theory, in 3d there is no local anomaly. However, there is a parity anomaly, due to the generation of Chern-Simons term upon integrating out massive charged fermions [4, 5, 6]. More explicitly, when integrating out charged fermions, the Chern-Simons term is generated at the one loop, and for the case of abelian gauge group (also abelian global symmetry group), the effective Chern-Simons levels in low energy become, 1 (kij )ef f = kij + (qf )i (qf )j sign(Mf ) 2 (2.15) where f labels fermions, i, j denote U (1)i gauge or global symmetries, (qf )i denotes the charge of f -th fermions charged under U (1)i symmetry, and Mf is the mass of f -th fermions. One can do similarly for the nonabelian global symmetry [2]. Due to the requirement that the path integral is invariant under large gauge transformation, the effective Chern-Simons level at the IR should be integer.1 2.1.4 Monopole operators and vortices Monopole operator [7, 8, 9] in 3d theory is defined as a prescription of singular behavior of gauge field; F ∼ qJ ? d 1 |~x − ~x0 | (2.16) In other words, insertion of a monopole operator at ~x0 creates a monopole source at that point with R flux qJ = S 2 F where S 2 encloses the insertion point ~x0 . 1 This integer condition is for dynamical gauge field. However, if we want to gauge the global symmetries, then the integer condition also needs to be imposed on them. 7 A chiral monopole operator in 3d N = 2 theories is expressed in terms of a linear mutliplet Σ; D2 Σ = 0, 2 D = qJ 2πδ (3) (~x − ~x0 )θ2 , (2.17) and similarly for anti-chiral monopole operator in obvious way. As seen in the definition, this is the UV operator, which is independent of the choice of the vacua. Though the monopole operator is a UV operator, it is related to (non-compact) Coulomb branch of the theory; Coulomb branch exists when the monopole operator is gauge neutral. In Coulomb branch, the linear multiplet Σ is massless and can be dualized to a chiral multiplet X, and the scalar component of X parametrize the (non-compact) Coulomb branch. X is charged +1 under U (1)J topological symmetry. Semi-classically, for large scalar component φ of the chiral multiplet Φ, the Coulomb branch modulus Xj where j labels the basis of simple roots βj of gauge group G is 2 asymptotically Xj ∼ eχ·βj /g where χ = φ + iγ with γ being a dual photon.2 This can be regarded as a semi-classical expression of a monopole operator. Meanwhile, there is a vortex solution in Higgs branch of the theory. For 3d N = 2 theories, one can have a BPS vortices under certain condition. For example, in the abelian Higgs model with an FI parameter ζ, there is a vortex configuration; φ ∼ ζ 1/2 e±θ , Aθ ∼ ± 1 r (2.18) with BPS mass given by m = |Z| = |qJ ζ|. As ζ → 0 where the Coulomb branch and Higgs branch meet, one can smoothly move onto Coulomb branch where U (1)J is explicitly broken by the Coulomb branch modulus charged under U (1)J . 2.1.5 3d N = 2 mirror symmetry There is a “mirror symmetry” in 3d N = 2 theories. Since there is no invariant way to distinct Higgs branch and Coulomb branch, which are exchanged under mirror symmetry, it is a misnomer. But since at least for known examples in SQED and SQCD they are from 3d N = 4 where there is actual mirror symmetry3 , same term is usually used also for 3d N = 2. The simplest example of 3d N = 2 mirror symmetry is the mirror symmetry between SQED with one flavor and XYZ model. This will be discussed in section 4 a bit more detail. 2 For abelian vector fields, it is possible to dualize it to scalar (dual photon) via F ρ µν = µνρ ∂ γ. Charge quantization condition constraints the dual photon γ to live on a circle whose radius is proportional to g 2 where g is a coupling constant of the gauge theory. The current J ρ = µνρ Fµν generates U (1)J symmetry which shifts γ by constant. 3 For example, mirror symmetry in 3d N = 4 theories whose R-symmetry group is SU (2) R1 × SU (2)R2 exchange SU (2)R1 and SU (2)R2 symmetries, Higgs branch and Coulomb branch, and mass term and FI term. 8 2.2 Compactification of 3d N = 2 gauge theories on S 1 In this section, we review reduction of 3d N = 2 theories on S 1 [10, 11, 12, 13]. Reduction of 3d N = 2 theories on S 1 is described by the effective 2d N = (2, 2) theories. This effective 2d supersymmetric theory is characterized by the twisted superpotential. Here, we only consider the case that all symmetries are abelian. Upon reducing 3d theory on S 1 with radius R, one can include a Wilson line for the (linear combination of) U (1) global symmetry and they complexity the (linear combination of) real mass parameter of 3d associated to the (linear combination of) global U (1) symmetry; mi = m3d i + i R I Ai (2.19) S1 where mi denotes (complexified) twisted mass or complexified FI term in 2d, which can be regarded as the scalar component of the background twisted chiral mutliplet Mi in 2d. Similarly, the scalar component σa of the 3d linear multiplet Σa are also complexified; σa = σa3d + i R I Aa (2.20) S1 where a runs from 1 to r, the rank of gauge group. If the abelian symmetry is compact, the invariance under large gauge transformation of 3d theory leads to the periodicity of σa and mi ; σa ∼ σa + 2πi R , mi ∼ mi + 2πi R . This is not an intrinsic property in 2d supersymmetric effective theory itself, but rather it is a property of the effective 2d supersymmetric theory obtained from the reduction of 3d theory. Upon reduction on S 1 , the chiral multiplet Φ in 3d gives a tower of Kaluza-Klein modes in 2d. If Φ is charged under overall U (1) whose overall real mass parameter is m3d Φ , the KK mode Φn with momentum n on S 1 has a twisted mass mΦn = mΦ + 2πin R . Due to the periodicity mentioned above, all KK modes should be included in the effective 2d theory. The twisted superpotential gets one-loop correction from integrating out the massive charged chiral multiplet. Taking into account all KK modes, one obtains [10, 11, 12, 13] f Φ) = 1 W(M R 1 2 MΦ + Li2 (−e−MΦ ) 4 (2.21) There are also contributions from tree-level mixed Chern-Simons terms, which is given by fCS (Σa , Mi ) = 1 W R 1 1 kab Σa Σb + kai Σa Mi + kij Mi Mj , 2 2 (2.22) 9 where in (2.21) and (2.22), we have absorbed R into MΦ and Σa to make them dimensionless.4 f the condition for supersymmetric vacua is given by Given the twisted superpotential W, f ∂W = 2πina , ∂σa na ∈ Z. (2.24) If one uses expontentiated variables valued in C∗ , sa = eσa and xi = emi , we get f ∂W exp sa ∂sa ! = 1. (2.25) If one gauged U (1)x symmetry associated to xi weakly and yi is the effective FI parameter for U (1)x , one can also consider the supersymmetric conditions for xi and yi , and they are given by f ∂W exp xi ∂xi ! = yi . (2.26) f is the twisted superpotential which was solved from (2.25), i.e. we solve (2.25) and put Here, W f and the resulting twisted superpotential is the one used in (2.26). We the solution sa back to W, will call the moduli space of (2.26) as the moduli space of supersymmetric vacua or supersymmetric parameter space. 2.3 The exact calculations in 3d N = 2 superconformal theories Since the pioneering work of Pestun [14], there has been many developments in exact calculations in various dimensions and with various operators or defects. In this section, we quote result from the recent development in exact calculations in 3d N = 2 theories [15, 16, 17, 18, 19, 20, 21]. 4 Expressions (2.21) and (2.22) are not single-valued under large gauge transformation, i.e. under the periodicity of σa and mi . In order to remedy this, one adds Z X exp −2πina d2 θΣa . (2.23) na ∈Z R into path integral, which ensures that Fa /2π ∈ Z that is required for the path integral to be well-defined. By adding this term, the resulting action becomes single-valued. 10 2.3.1 Supersymmetric partition function on the squashed 3-sphere Our main interest is the partition function on squashed 3-sphere Sb3 preserving U (1)×U (1) symmetry [18]. Metric for such squashed 3-sphere Sb3 is ds2 = b2 (dx20 + dx21 ) + b−2 (dx22 + dx23 ), with x20 + x21 + x22 + x23 = 1 (2.27) The partition function on Sb3 takes the form Z Sb3 1 = |W| Z dr σ Zclassical Zone-loop (2.28) where |W| is the order of Weyl group and r is the rank of gauge group. If chiral multiplet has R-charge R and gauge charge ρa under the Cartan subgroup Ha of gauge group G, then the one-loop determinant from chiral multiplet is given by sb iQ (1 − R) − ρa σa 2 (2.29) where σa is (scaled) scalar component of vector multiplet and Q = b + b−1 . sb (x) is sine double function and is given by sb (x) = Y m,n∈Z≥0 mb + nb−1 + Q 2 − ix mb + nb−1 + Q 2 + ix iπ = e− 2 x 2 ∞ Y j=1 2 1 1 + e2πbx+2πib (j− 2 ) −1 x+2πib−2 ( 1 −j) 2 1 + e2πb (2.30) which is a variant of non-compact quantum dilogarithm function. One-loop determinant from vector multiplet is trivial for abelian gauge group, but for nonabelian gauge group G it is given by Y sinh(πbσα) sinh(πb−1 σα) πσα (2.31) α∈∆+ Here, α takes positive roots and σα denotes σα = Pr a=1 σa αa . There are non-zero contributions from classical action of Chern-Simons term and FI term at the saddle point. For Chern-Simons term with level k and FI term for U (1) gauge group with FI parameter ζ, they are given by exp(−ikπσa σa ) (2.32) 11 and exp(4πiζσ), (2.33) respectively. Assuming that R-symmetry at the IR fixed point is not mixed with the accidental symmetry, it is 3 known that the IR superconformal R-charges are fixed by maximizing F = − log |ZSb=1 | [16, 22, 23]. 2.3.2 Superconformal index Superconformal index counts certain BPS operators in superconformal field theory on S 2 ×q S 1 [19, 20, 21]. More explicitly it is given by Y F βH ∆+j3 i I(ti ; q) = tr (−1) e q tF i (2.34) a where ∆ is energy, R is R-charge, j3 is angular momentum on S 2 , Fi are generator for the global (or flavor) symmetry, and H is given by H = {Q, Q† } = ∆ − R − j3 (2.35) where Q is a certain supercharge in 3d N = 2 superconformal algebra. Since only Q and Q† -invariant states contribute to the index, index (2.34) doesn’t depend on β, but only depends on fugacities, q, ti . From the localization technique, one can obtain the exact result for the index. But if one incorporates the magnetic fluxes and Wilson line for flavor symmetries, then the resulting index has extra discrete and continuous parameters. This is called the generalized index. It is given by I(ti , ni ; q) = X s 1 Sym Z Y a Y dza −SCS (h,m) e Zgauge (za , ma ; q) ZΦ (za , ma ; ti , ni ; q) 2πza (2.36) Φ where Zgauge (za = eiha , ma ; q) = Y α∈∆+ 1 q − 2 |α(m)| 1 − eiα(h) q |α(m)| (2.37) 1 Y 1 Y Y −f (Φ) 2 |ρ(s)+ (1−∆Φ ) −iρ(h) iha 2 ZΦ (za = e , ma ; ti , ni , q) = q e tj j a ρ∈RΦ × q f (Φ) (eiρ(h) tj j q |ρ(s)+ j fj (Φ)nj | (2.38) j P −f (Φ) |ρ(s)+ j fj (Φ)nj |+1− ∆Φ 2 (e−iρ(h) tj j P ∆Φ j fj (Φ)nj |+ 2 P ; q)∞ ; q)∞ (2.39) 12 and Sym is a symmetric factor arising from non-abelian gauge group generically broken by monopole, Q i.e., if gauge group G is broken to its subgroup ⊗k Gk by monopole then Sym = k=1 Rank(Gk )!. Here, α is positive roots of gauge group G, h whose component is ha denotes maximal torus of the gauge group which parametrize the Wilson line of the gauge group, and s whose component is ma denotes the Cartan generator of the gauge group which parametrize the GNO charge of magnetic monopole configuration of gauge field. a runs over the rank of the gauge group, and takes integer value. If chiral multiplet Φ is in representation R of gauge group, ρ denotes a weight vector of representation R, R denotes R-charge of chiral multiplet, tj denotes the maximal torus of the global symmetry under which chiral multiplets Φ are charged, nj denotes the Cartan generator of the global symmetry group which parametrize the GNO charge of magnetic monopole configuration of background gauge field for global symmetry, and fj (Φ) denotes the charge of chiral multiplet Φ under U (1) subgroup of maximal torus associated to tj . j runs over the rank of the global symmetry group, and takes integer value. Contribution from Chern-Simons term is given by eikSCS (h,s) = exp(ik P a ha ma ) = kma . a za Q For the mixed Chern-Simons term with level k12 = k21 between two abelian symmetries U (1) with k12 Wilson line and magnetic flux being (h1 , m1 ) and (h2 , m2 ) is given by z1m2 z2m1 . 2.4 Holomorphic block in 3d N = 2 theories It turns out that above partition functions on squashed 3-sphere Sb3 − ZSb3 − and superconformal index on S 2 ×q S 1 − IS 2 ×q S 1 − are factorized [24, 25], and this was systematically studied in [13]. The basic ingredients for ZSb3 and IS 2 ×q S 1 in 3d N = 2 theory which have discrete, massive, supersymmetric vacua (i.e. there should be enough flavor symmetries so that it completely lifts all flat directions of the moduli space) are called as holomorphic block. Holomorphic block is defined as a partition function on D2 ×q S 1 which is topologically a solid torus and whose metric is given by ds2 = dr2 + f (r)2 (dϕ + βdθ)2 + β 2 dθ2 (2.40) where local coordinate is (r, ϕ, θ) with r ∈ [0, ∞), ϕ ∼ ϕ + 2π, and θ ∼ θ + 2π. f (r) approaches to 0 as r → 0 and ρ as r → ∞. The geometry D2 ×q S 1 can be regarded as cigar geometry parametrized by (r, ϕ) is fibered over S 1 parametrized by θ in such a way that (z, θ) ∼ (q −1 z, θ + 2π) if we set z = reiϕ and q = e2πiβ = e~ . (2.41) 13 Since it is curved, in order to preserve supersymmetry, appropriate twisting is necessary. There are two choices of twisting, one is “topological” and another is “anti-topological”. They preserve two different sets of two supercharges with different BPS conditions (P 0 = Z for the former and P 0 = −Z for the latter). 2.4.1 Holomorphic block as the BPS index Since the theory is twisted, partition function doesn’t depend on the size of ρ. So one can equivalently think of the theory on R2 ×q S 1 . In this setup, one can associate supersymmetric state at the origin of R2 and also supersymmetric state |αi on T 2 at infinity (since asymptotically R2 ×q S 1 approaches to T2 × R) which one-to-one corresponds to discrete, massive, supersymmetric vacua α of 3d N = 2 theories. Roughly speaking, their overlap give the BPS index. If we deform superconformal field theory away from the superconformal fixed point, we can obtain massive theory, and the BPS index counts BPS states in the vacuum of the massive theory. Given this setup, one can consider holomorphic block as BPS index for each vacua α; α B (x; q) ≃ ( α ZBPS (x; q) |q| < 1 (2.42) α ZBPS (x; q) |q| > 1 with R α ZBPS (x; q) = TrHα (−1)R e−βH q −J3 + 2 x−e R e α (e x; qe) = TrHα (−1)R e−β H qe−J3 − 2 x ZBPS e−e |q| < 1 (anti-topological) |e q| > 1 (topological) (2.43) e = 2(P 0 − Z), and x and x where H = 2(P 0 + Z), H e are certain exponential of complexfied twisted mass parameter associated to U (1)x global symmetry. In the BPS indices, (−1)R was chosen instead of usual choice (−1)F , but both give essentially equivalent information. As varying x, there is an interesting stoke phenomena, which we don’t discuss here. Holomorphic 2 1 block is unique modulo the overall factor of an elliptic ratio of theta functions or exp − 24 (~ − 4π~ ) . 2.4.2 Fusion of holomorphic and anti-holomorphic block In analogue of topological/anti-topological fusion in 2d N = 2 theories [26], one can consider fusion of two copies of D2 ×q S 1 . Fusion is possible when Hilbert spaces on the each asymptotic regions are dual [13]. Considering two infinitely long cigar geometries where D2 ×q S 1 is asymptotically T 2 × R, we would like to perform fusion by gluing the two tori with opposite orientation via SL(2, Z) action on the torus, which also induces modular action (combined with reflection) on the parameter q. A bit more specifically, we would like to consider fusion such that if the complex structure 14 τ = β + iβρ−1 of T 2 in infinitely long cigar geometry for anti-holomorphic block is related to an infinitely long cigar geometry with complex structure τe by τ → τe = −g · τ for holomorphic block e = −g · (β). then β → βe Then for S-fusion one obtains q = exp(2πib2 ), qe = exp(2πib−1 ) and x =: exp(X) = exp(2πbµ), e = exp(2πb−1 µ) where b is squashing parameter of squashed 3-sphere S 3 mentioned x e =: exp(X) b above and µ are complexified mass parameter. For id-fusion one obtains q = qe−1 and x =: exp(X) = m m e = q 2 ζ −1 where q = exp(2πβ) and m and ζ are magnetic flux and fugacity of the q 2 ζ, x e =: exp(X) index associated to U (1) global symmetry. Along the way of obtaining these, in order to have exact same holomorphic block for both Sb3 partition function and the index, R-charges of chiral multiplets need to be all integers. In sum, via fusion, one obtains [g] Zfusion = [S] X 2 B α (x; q)B α (e x, qe) = B α (x; q) g (2.44) α [id] where we are interested Zfusion or Zfusion , which give the partition function on Sb3 and the index on S 2 ×q S 1 . 2.4.3 Holomorphic block from supersymmetric quantum mechanics The cigar geometry D2 ×q S 1 can be considered as torus fiberation over half-line R+ = {t ∈ [0, ∞)} with fixed area and complex structure of torus as t → ∞. Thus, macroscopically 3d N = 2 theory can be seen as supersymmetric quantum mechanics on R+ . From the point of view of supersymmetric quantum mechanics, one can see that path integral on D2 ×q S 1 becomes a finite-dimensional contour integral with the integrand determined perturbatively all order of ~ [13]. The result is that for a discrete, massive, supersymmetric vacua labelled by α, there is an associated integration cycle Γα , and contour integral with integrand from twisted superpotential gives holomorphic block B α . A bit more specifically, it turns out that the path integral of supersymmetric quantum mechanics become a contour integral B α α (x; q) = ZQM ≃ ds1 dsr ··· exp 2πisr Γα 2πis1 Z 1f W~ (sa , xi ; ~) ~ (2.45) in complex space C∗ r parametrized by sa where we use the notations in section 2.2. The integration cycle Γα should be convergent and determined by gradient flow equation with respect to ImW QM given a choice of asymptotic boundary condition at t → ∞ labelled by the massive vacua α. Here, 15 superpotential W QM of quantum mechanics is given by W QM f a , Mi ) = i (Σa , Mi ) = 2πρW(Σ ~ " X 1 Φ 4 MΦ2 + Li2 (−e−MΦ ) 1 1 + kab Σa Σb + kai Σa Mi + kij Mi Mj 2 2 (2.46) which is obtained from further compactification on Sρ1 of the twisted superpotential obtained in f~ (sa , mi ; ~) in the integrand is the “quantumsection 2.2 modulo a certain subtlety. The term W corrected superpotential” f~ (sa , xi ; ~) = W where Li2 (x; ~) = ~ 1 1 m2Φ + Li2 (−e−mΦ − 2 ; ~) + kab σa σb + kai σa mi + kij mi mj 4 2 2 X 1 Φ P∞ n=0 Bn ~ n n! (2.47) with Bernoulli number Bn . The R-charge of chiral multiplet is set to one here. It is also possible to have general R-charge RΦ by shifting, mΦ → mΦ +(RΦ −1)(iπ +~/2). f~ (sa , xi ; ~) becomes ~ W QM (σa , mi ). If we take ~ → 0 limit, quantum-corrected superpotential W i Though above analysis from supersymmetric quantum mechanics gives rough idea that the holomorphic block is expressed as certain finite-dimensional contour integral, as stated earlier this is a perturbative analysis for both integrand and the integration cycle. With help of Ward identity of Wilson and ’t Hooft loop operators, the exact non-perturbative integrand can be obtained. However, since the exact superpotential whose only critical points correspond to the true vacua of the theory5 , which is used in gradient flow equation to determine exact Γα , is still unknown, one can just use approximated Γα from above perturbative argument [13]. With several conditions imposed on the Γα in perturbative analysis, one can obtain exact non-perturbative holomorphic block. Those conditions are that if one shifts entire Γα by q, one should be able to deform smoothly it back to original Γ before shift. This means that the contour should be closed or end asymptotically at 0 or ∞ in each C∗ r . Also, Γα should be far away at least by q from the poles of the integrand. In addition, Γα is not allowed to path line of poles but is allowed to cross or lie on the line of zeroes. 2.5 Sp(2L, Z) action on 3d conformal field theories For 3d conformal field theories with U (1) symmetry which one can couple to U (1) background  gauge 0 symmetry, there is an SL(2, Z) action on them. This action is generated by S =  −1   1 1 . T = 1 0 1  and 0 5 A non-perturbative integrand Υ obtained with help of Ward identity of line operators has too many critical points including the true vacua. # 16 The generator T acts on the theory in a way that it adds Chern-Simons coupling with level 1 for background gauge field A; L→L+ 1 A ∧ dA. 4π (2.48) The generator S acts on the theory in a way that it gauges U (1) symmetry and introduces a new background U (1)new gauge field Anew by coupling Anew with the current that is a Hodge star of field strength for U (1) symmetry which is now gauged. In terms of Lagrangian, we add a mixed ChernSimons term of U (1) symmetry which is now gauged and newly introduced U (1)new background gauge symmetry; 1 Anew ∧ dA 2π L→L+ (2.49) where A is dynamical gauge field of U (1) symmetry which is now gauged. A new U (1) background gauge symmetry whose gauge field is Anew is usually called as topological symmetry with respect to A, because monopole operator defined from A is charged under this new U (1) symmetry. When 3d conformal field theory is coupled to 4d free abelian theory as a boundary theory, this SL(2, Z) action can be thought as electro-magnetic duality on a space of conformally invariant boundary condition of a 4d free abelian theory. If there are more than one U (1) symmetry, say L U (1) global symmetries, SL(2, Z) action can be generalized to Sp(2L, Z) action on the 3d conformal field theory with U (1)L global symmetries. ~ = (A1 , A2 , · · · , AL ) be gauge fields for U (1)L global symmetries. For Sp(2L, Z) action, Let A there are three types of generators, which are called as T -type, S-type, and GL-type. More explicitly, by representing them as L × L block matrices, they are expressed and act on Lagrangian as follows;  T -type I gT =  B 0 I  : 1 ~ ~ new Anew · B dA 4π (2.50) ~ → L[A] ~ + 1 A ~ new · J dA ~ L[A] 2π (2.51) ~ → L[A ~ new ] + L[A] where B is a symmetric L by L matrices;  S-type gS =  I −J −J J I −J  : where J is a diagonal matrix J = diag(j1 , j2 , · · · , jL ) with ji is 0 or 1. When ji = 1, we gauge U (1) symmetry whose gauge field is Ai and introduce U (1) topological symmetry whose gauge field is 17 Anew,i ;  U gU =  0 GL-type 0 U −1 t  : ~ → L[U −1 A ~ new ] L[A] (2.52) where U ∈ GL(N, Z) is invertible, which redefines global symmetries. Above argument can also be extended to supersymmetric case. More explicitly, for vector mul~ and linear multiplet Σ ~ we have tiplet V  T -type S-type GL-type I 0  Z ~ ] → L[V ~new ] + 1 ~ new · B dV ~new  : L[V d4 θ Σ 4π B I   Z I −J −J 1 ~ new · J dV ~ ~ ~   d4 θ Σ gS = : L[V ] → L[Vnew ] + 2π J I −J   I 0 ~ ] → L[U −1 V ~new ]  : L[V gU =  B I gT =  (2.53) (2.54) (2.55) We will see in Chapter 4 that how Sp(2L, Z) will be related to the operation in Chern-Simons theory. 18 Chapter 3 Chern-Simons Theory In this chapter, we review several aspects of Chern-Simons theory. We follow closely [27, 28, 29, 30, 31, 32]. Before we review analytically continued Chern-Simons theory, we briefly summarize Chern-Simons theory which is not analytically continued. When the gauge group is G with Lie algebra g, gauge field A is a connection on G-bundle E on 3-manifold M3 , and Chern-Simons action is given by I(A) = k 4π 2 Tr A ∧ dA + A ∧ A ∧ A 3 M3 Z (3.1) where Tr is a trace on fundamental representation of the gauge group. If we take a proper normalization for the gauge group generator Tr(Ta Tb ) = δab , due to the requirement of invariance of the path integral Z(M3 ) = Z DA exp(iI(A)) (3.2) under large gauge transformation, Chern-Simons level k should be an integer, k ∈ Z. The observable in Chern-Simons theory is Wilson loop operator which depends on the representation of gauge group and is supported on a knotted curve (or more generally link) in M3 . In the presence of Wilson loop operators, the partition function is Z(M3 ; γi , Ri ) = Z DA exp(iI(A)) r Y WRi (γi ) (3.3) i=1 where Wilson loop operator WRi (γi ) is given by I WRi (γi ) = TrRi P exp i A . γi (3.4) 19 where Ri is representation of gauge group G for each i-th knot in link. It turns out that Wilson loop operator expectation value gives Jones polynomial of knot γ in M3 . As a simple example, for unknot (01 ) in S 3 , the expectation value of Wilson loop operator in r-symmetric representation of G = SU (2) is given by r hWS r (01 )i = 2 sin k+2 (r + 1)π 2 , (3.5) which agrees with Jones polynomial of unknot. Expectation value of Wilson loop operator on more complicated knots can be calculated via surgery of 3-manifold by calculating corresponding matrix elements in WZW model where the Hilbert space of Chern-Simons theory is equivalent to the space of conformal block of WZW model whose level is k [31]. Later, Chern-Simons theory with complex gauge group was investigated [33], and analytic continuation of Chern-Simons theory with complex gauge group was studied in [27] and also in [30, 28, 29]. These and further developments will be a main review topic of this chapter. In section 3.1, we review the structure of analytically continued Chern-Simons theory with complex gauge group GC and its real form GR and G. Mostly, we take G = SU (2), so GR = SL(2, R) and GC = SL(2, C). We also briefly mention some aspects of Wilson loop. In section 3.2, we review classical moduli space of Chern-Simons theory on a knot complement, which is known as A-polynomial. This will be related to the moduli space of supersymmetric vacua or supersymmetric parameter space of effective 3d N = 2 theories on S 1 corresponding to knot b complement. Also, we discuss briefly quantum A-polynomial, which gives recursion relation for knot polynomials. In section 3.3, we review some aspects of Chern-Simons theory on 3-manifold obtained from gluing tetrahedra. An ideal tetrahedron is a basic building block of 3d-3d correspondence of DimofteGaiotto-Gukov, so we review some properties of an ideal tetrahedron and gluing procedure for classical moduli space, quantum operator, and partition function. 3.1 Analytic continuation of Chern-Simons theory Given gauge group GC with Lie algebra gC , let the gauge field A be a connection on GC bundle EC on three manifold M3 . We first consider a complex Chern-Simons action, I(t, e t) = t 8π e t 2 Tr(A ∧ dA + A ∧ A ∧ A) + 3 8π M3 Z 2 Tr(A ∧ dA + A ∧ A ∧ A) 3 M3 Z (3.6) 20 where A is a complex conjugate of A and in order for it to be real t and e t should be complex conjugate. If we denote t = k + is and e t = k − is and require invariance of the path integral under the large gauge transformation then k should be integer k ∈ Z. In addition, unitarity requires s to be real or imaginary for Euclidean action [33]. From the above action, we can analytically continue both t and e t and regard them separate and independent variables. At the same time, we also analytically continue A. This leads to (gC )C = gC ⊕ gC . So we have two copies of gauge connection, which are regarded as independent. e Then the action for analytically continued complex Chern-Simons We denote them as A and A. theory is given by e t, e IC (A, A; t) = t 8π e 2 t Tr(A ∧ dA + A ∧ A ∧ A) + 3 8π M3 Z 2 e Tr(Ae ∧ dAe + Ae ∧ Ae ∧ A) 3 M3 Z (3.7) So we can think of analytically continued complex Chern-Simons theory as sum of two independent Chern-Simons theory whose levels are k1 = 2t and k2 = 2t . e We can also consider the case of compact group G and non-compact real group GR . From (3.1), we analytically continue level k to arbitrary complex number t. At the same time, we also analytically continue the gauge field A to GC gauge field A. The resulting action becomes I(A; t) = t 8π Z M3 2 TrA ∧ dA + A ∧ A ∧ A 3 (3.8) where we took 1/2 normalization to the action. We can regard (3.8) as “holomorphic half” of (3.7). Similar argument is also applied to the case of non-compact real form GR of GC . 3.1.1 Structure of Chern-Simons partition functions The partition function of Chern-Simons theory is obtained from path integral ZC (M3 ) = Z C e DADAe exp(iIC (A, A)). (3.9) If s is not real, the path integral is generically not convergent, so it is desirable to find a welldefined integration cycle such that the path integral is well-defined. In addition, when s is real such integration cycle should be the middle dimensional cycle such that Ae = A. In [30], such integration cycle C was analyzed via Morse theory and the steepest descent method. Briefly summarizing, integration cycle is associated to each set of critical points of the action 21 e The critical points Aα of the action are given by flat GC connections on M3 , I(A, A). F = Fe = 0, (3.10) e It is known that there are finite number of critical where F = dA + A ∧ A and Fe = dAe + Ae ∧ A. points in many 3-manifolds including knot complement in 3-manifold with appropriate boundary conditions along the knot complement and without closed incompressible surface [34, 29]. Given a critical point, integration cycle is given by a downward gradient flow from it with Morse function from GC Chern-Simons functional. This integration cycle Jα associated to a critical point Aα is called Lefschetz thimbles modulo some subtleties. Then the convergent integration cycle C is given P by C = α nα Jα where nα ∈ Z, which is calculated in Morse theory. Since Morse function depends on GC Chern-Simons functional which depends on s, Lefschetz thimble Jα depends on the value of s. Similar argument applies for the analytic continuation of Chern-Simons theory for compact real form G and also for non-compact real form GR . The integration cycle Ccompact for analytically P continued Chern-Simons theory for compact group G is given by Ccompact = α nα Jα . For the 0 non-compact real form GR , the integration cycle C is also given by linear combination of Jα over Z for each critical point which corresponds to critical point Aα . In sum, for any GC , GR , and G, the downward flow cycle Jα corresponds to critical point Aα of 0 flat GC connection and they form a vector space over Z. But the integer coefficient nαe α , nα , and nα depends on a specific integration cycle in a particular Chern-Simons theory under consideration. Thus, given the integration cycle Jα , the form of partition function of SL(2, C), SL(2, R), and SU (2) Chern-Simons theory has the following form ZSU (2) (M3 ) = X ZSL(2,R) (M3 ) = X ZSL(2,C) (M3 ) = X nα Zα (M3 ) (3.11) n0α Zα (M3 ) (3.12) ee (M3 ) nα,e α Zα (M3 )Zα (3.13) α α α,e α where Z Zα (M3 ) = exp(iI(A)). (3.14) Jα In (3.13), nα,e α is diagonal for non-anlytically-continued Chern-Simons theory where integration e but in general it can be arbitrary. cycles satisfy A = A, 22 3.1.2 Wilson loop operator Wilson loop operator is an observable in Chern-Simons theory. It is labeled by a knotted loop on which the Wilson loop is supported and the representation of gauge group. We can also have Wilson loop in analytically continued Chern-Simons theory, I WR (A; γ) = TrR P exp i A (3.15) γ There are two ways to think of Wilson loop operator. We can regard it as a path-ordered integral of gauge field on 1-dimensional curve γ as above. However, we can also consider it as a particle moving around γ. In this case, the particle creates defect of cusp along the trajectory and inserting Wilson loop operator in 3-manifold can be interpreted as giving a boundary condition for gauge field so that one has holonomy around γ. In other words, we can consider it as ’t Hooft operator, which we also call as monodromy defect. Actually, in 3d, it was shown that ’t Hooft operator is equivalent to Wilson operator [32]. For example, for r-dimensional irreducible representation of SU (2), the monodromy defect gives the boundary condition for gauge field such that it has holomony  eπir/k  0 0 e−πir/k   (3.16) where k is a renormalized level here. This alternative interpretation is useful. For compact gauge group, say SU (2), we can use both interpretation interchangeably. However, considering the infinite dimensional representation of complex gauge group, for example SL(2, C), it is more natural to consider monodromy defect.1 Detail discussion on infinite dimensional representation is summarized in Chapter 7 of [35]. 3.2 b Classical moduli space, A-polynomial and quantum Apolynomial of knot In this section, we review the classical vacuum moduli space of Chern-Simons theory on a knot complement in three sphere S 3 , and quantization of it. 3.2.1 Classical moduli space and A-polynomial Consider knot complement in 3-manifold, M3 , for example, S 3 \K where K is a knot. The boundary of a knot is a torus, T 2 , so there are two curves γm and γl corresponding to meridian and longitude curve, respectively. 1 For example, it is unclear what would be Tr in above equation for infinite dimensional representation. 23 K Figure 3.1: Meridian cycle γm and longitude cycle γl on the boundary of a knot complement Holonomies around γm and γl generate the peripheral subgroup of M3 , which is π1 (T 2 ) = Z × Z. So SL(2, C) holonomies around γm and γl can be simultaneously taken in upper triangular matrices  m ρ γm =  0 ∗ m−1   , ρ γl =  l 0 ∗ l−1   (3.17) with eigenvalues l, m ∈ C∗ . The action of Weyl group Z2 of SL(2, C) on (3.17) exchanges (l, m) and (l−1 , m−1 ). Thus, the classical phase space for M3 with a single boundary given by a knot is PT 2 = (C∗ × C∗ )/Z2 (3.18) We are interested in the moduli space of flat connection, and it is given by the homomorphism from fundamental group of M3 to SL(2, C), Mflat (SL(2, C), M3 ) = Hom(π1 (M3 ), SL(2, C))/conj. =: L (3.19) It is known that for a single torus boundary of M3 , the complex dimension of L is 1. Two generators ργm and ργl of peripheral subgroup determines embedding of L into PT 2 , and it can be shown that L is characterized by the zero locus of a single polynomial, L = {(l, m) ∈ PT 2 |A(l, m) = 0} (3.20) This also can be thought as the condition that the flat connection on T 2 is extended to the flat connection on the bulk M3 . This polynomial A(l, m) is called as A-polynomial of a knot [34]. For every knot complement, there is always (l − 1) factor. This corresponds to the abelian sector of L where SL(2, C) flat connection are simultaneously diagonalized. Since H1 (M3 , Z) = Z for any knot complement, there is always abelian representation in L which is generated by the holonomy around meridian curve and holonomy around longitude curve is trivial implying that l − 1 = 0. 24 As an example, A-polynomial of unknot, trefoil knot, and figure-eight knot are A01 (l, m) = (l − 1) (3.21) A31 (l, m) = (l − 1)(l + m6 ) (3.22) A41 (l, m) = (l − 1)(l2 − (m4 + m−4 − m2 − m−2 − 2)l + 1) (3.23) In classical phase space and also in quantum phase space, we can only specify one of l or m, but not both. In the following, we will specify m. For a given m, i.e. the boundary condition on the holonomy of flat SL(2, C) connection around meridian cycle on the torus boundary, the order of l in A-polynomial is then the number of flat connection. 3.2.2 b Quantization and quantum A-polynomial In analytically continued Chern-Simons theory, the classical phase space is given by PT 2 = (C∗ × C∗ )/Z2 and (semi)-classical state is given by algebraic curve L ⊂ PT 2 . Considering the Hamiltonian approach in [33], one can quantize analytically continued Chern-Simons theory [27]. The holomorphic symplectic structure on the (semi)-classical phase space PT 2 is induced from the holomorphic Chern-Simons term ωT 2 = 2 d log l ∧ d log m, ~ (3.24) 2 2 where ~ = 2πi k . If we introduce logarithmic variables v = log l, u = log m for l and m, ωT = ~ dv∧du. It can be shown that the algebraic variety L given by {A(l, m) = 0} is a Lagrangian submanifold in phase space (PT 2 , ωT 2 ). Also from ωT 2 , commutation relation is given by [b v, u b] = ~/2, and this leads to b lm b = q 1/2 m bb l, (3.25) where b l = evb, m b = eub , and q = e~ . Correspondingly, under certain conditions, classical A-polynomial b can be quantized to quantum A-polynomial [27, 29, 36, 37, 38], A(l, m) bb A( l, m; b q) (3.26) Upon quantization, the classical state L = {A(l, m) = 0} becomes the wavefunction in the Hilbert space HT 2 . A wavefunction in Hilbert space is a Chern-Simons partition function (3.14), and this 25 bb is annihilated by A( l, m; b q). u b and vb act on the wave function f (u); vbf (u) = ~ ∂u f (u), 2 u bf (u) = uf (u). (3.27) b bb As classical A-polynomial A(l, m) always have (l − 1) factor, quantum A-polynomial A( l, m; b q) also always have (b l − 1) factor. 3.2.3 Recursion relation for Jones polynomial Above discussion leads to the recursion relation of Jones polynomial. A colored Jones polynomial Jr (K; q) is obtained from SU (2) Chern-Simons partition function on knot complement in S 3 and r denotes the dimension of the irreducible representation of SU (2). Then the holonomy around the meridian cycle is conjugated to (3.16). So we have appropriate change of variables; u = iπr/k with ~ = 2πi/k. Thus, u bJr (K; q) = q r/2 Jr (K; q), vbJr (K; q) = Jr+1 (K; q) (3.28) In general, full Chern-Simons partition function such as Jones polynomial is given by linear combination of partition function as in (3.11) for both abelian and non-abelian flat SL(2, C) connections. In other words, Jr (K; q) = X nα Zα (3.29) α where α contains both abelian and non-abelian holonomy of SL(2, C) flat connections. Upon quanb tization, this full partition function is annihilated by proper quantum A-polynomial which always include (b l − 1) factor in the left of everything else, and there are also nontrivial additional factor corresponding to non-abelian flat SL(2, C) connections. If we take off (b l − 1) part and call the rest b bnab , A bnab cannot annihilate the full partition function. More of full quantum A-polynomial as A explicitly, for Jones polynomial, Anab (b l, m)J b r (K; q) = B(m; q) 6= 0 (3.30) After diving by B(m; q) for both sides, multiplying (b l − 1) gives RHS to be zero. After some calculation, we obtain bnab (b (B(m; b q)b l − B(q 1/2 m; b q))A l, m; b q) Jr (K; q) = 0, (3.31) b bnab for non-abelian and the LHS is what we mean by full quantum A-polynomial. For example, A 26 flat connections for trefoil knot, and figure-eight knot in S 3 are b b q) = b bnab A l + q 3/2 m b6 31 (l, m; (3.32) b b q) = q 5/2 (1 − q m bnab A b 4 )m b 4b l2 − (1 − q 2 m b 2 )(1 − q m b − (q + q 3 )m b 4 − q3 m b 6 + q4 m b 8 )b l + q 3/2 (1 − q 3 m b 4 )m b4 41 (l, m; (3.33) bnab (b By taking q → 1 (or ~ → 0), above A l, m; b q) becomes classical A-polynomial for non-abelian flat connection. There are commutative deformations of classical and quantum A-polynomials. Such quantum b A-polynomials annihilate or give recursion relation for the corresponding Poincaré polynomials of the sl(N ) knot homology or the HOMFLY homology. This will be discussed in Chapter 5. For above two examples, it can be obtained via gluing ideal tetrahedra, which is a next topic. 3.3 Ideal hyperbolic tetrahedra and their gluing In [29], Chern-Simons theory on 3-manifolds M3 which admits tetrahedra triangulation were considered. The purpose of this section is to give a general idea on ideal tetrahedra and gluing them. Detail construction and relevant subtleties are explained in [29, 39, 28]. In this section, we only consider g = su(2). 3.3.1 An ideal tetrahedron, boundary phase space, and Lagrangian submanifold It is known that one can describe flat SL(2, C) structure in terms of hyperbolic geometry.           Figure 3.2: An ideal hyperbolic tetrahedron A hyperbolic 3-manifold H3 is a 3-manifold if it admits a hyperbolic metric, which is a metric 27 whose constant curvature −1, of finite volume. A hyperbolic 3-manifold can be regarded as inside of 3-ball B 3 with boundary ∂H3 = C ∪ {∞} which is a Riemann sphere. An ideal hyperbolic tetrahedron is a tetrahedron whose vertices are on the boundary of hyperbolic space ∂H3 and whose faces are geodesic surface in H3 . Its full hyperbolic structure is determined by a single cross ratio of the position of the vertices on ∂H3 , which is called the shape parameter. By using isometry group P SL(2, C) of hyperbolic 3-space which acts on the boundary as Möbius transformation, it is possible to fix three vertices at 0, 1, and ∞. Let the position of remaining vertex be z ∈ C∗ \{1}. Then the opposite edge of the ideal tetrahedron has dihedral angle arg(z), so we label those two edges as z. Let other two edge parameters as z 0 and z 00 . These edge parameters are related to z; zz 0 z 00 = −1 (3.34) z + z 0−1 − 1 = 0 (3.35) z 0 + z 00−1 − 1 = 0 (3.36) z 00 + z −1 − 1 = 0 (3.37) and any one of three equivalent forms; where we will choose the last one. Or in logarithmic variables, Z = log z and similarly for z 0 and z 00 , we have Z + Z 0 + Z 00 = iπ, 00 eZ + e−Z − 1 = 0. (3.38) The boundary phase space is given by P∂∆ = {(Z, Z 0 , Z 00 ) ∈ (C\2πiZ)3 |Z + Z 0 + Z 00 = iπ}, (3.39) and this affine linear space is equipped with symplectic form ω∂∆ = 1 00 dZ ∧ dZ ~ (3.40) where ~ is chosen as a normalization for the symplectic form. Or equivalently we have Poisson brackets; {Z, Z 0 } = {Z 0 , Z 00 } = {Z 00 , Z} = ~. (3.41) A polarization Π∂∆ for the boundary phase space P∂∆ is about what we choose for position and 28 conjugate momentum with respect to symplectic form ω∂∆ . For a tetrahedron, we have three sets of position X and conjugate momentum P ; (X, P ) : ΠZ = (Z, Z 00 ), Π0Z = (Z 0 , Z), Π00Z = (Z 00 , Z 0 ). (3.42) They are related by the action of affine symplectic transformation Sp(2, Z)n(iπZ)2 . More explicitly, for example, from the relation above, one can map ΠZ to Π0Z via    Z0 −1  = Z 1   iπ   +  . 0 Z 00 0 −1  Z  (3.43)     1 1 1  and T =  . Above matrix is expressed as ST ∈ Sp(2, Z) ≃ SL(2, Z) where S =  −1 0 0 1 Since (ST )3 = I, one polarization gets back the original polarization after the action of ST three 0 times. Lagrangian submanifold of P∂∆ of a tetrahedron is 00 L∆ = {eZ + e−Z − 1 = 0} ⊂ P∂∆ . (3.44) We have seen a boundary phase space and Lagrangian submanifold of a single tetrahedron, and we would like to see what they become upon gluing a number of tetrahedra. 3.3.2 Gluing tetrahedra and boundaries of 3-manifold Before actually mentioning the gluing procedure, we consider the boundaries of 3-manifold from gluing tetrahedra. If a 3-manifold M3 obtained by gluing ideal tetrahedra, there are two kinds of boundaries, which are geodesic boundary and cusp boundary. The geodesic boundary is a geodesic surface of any genus possibly with punctures, and there is an induced 2d hyperbolic metric on it. 3d triangulation determines 2d triangulation on the geodesic boundary. This comes from the unglued faces of tetrahedra. The cusp boundaries don’t allow such triangulation. They are knotted loci in M3 along which hyperbolic metric has a cone angle, or around which the flat SL(2, C) connection has a monodromy. If one replaces ideal tetrahedra participating in gluing by the ideal tetrahedra whose four vertices are truncated or regularized, the loci are resolved to the boundary of M3 with topology of tori T 2 or annuli S 1 × I where I denotes interval. The latter connects the punctures on the geodesic boundary of M3 . An induced metric on those boundaries are Euclidean. 29 Going back to gluing, we want to see how we get the boundary phase space P∂M3 and Lagrangian submanifold LM3 for M3 obtained from gluing ideal tetrahedra. Given a collection of a number of tetrahedra, what we mean by gluing is that the faces of tetrahedra are glued such that three edges of the each faces are also glued. If we want to make hyperbolic metric on M3 to be smooth on the internal edge from the gluing, the total dihedral angles around the internal edge should be 2π with zero hyperbolic torsion. More explicitly, if CI is the internal edge from the gluing, the following should hold CI = L X (n(I, i)Zi + n0 (I, i)Zi0 + n00 (I, i)Zi00 ) = 2πi (3.45) i=1 where L is the total number of tetrahedra and n(I, i) denotes how many edges are glued to I-th internal edge. So n(I, i) takes value 0,1,2,2 and similar for n0 (I, i) and n00 (I, i). We are interested in the boundary phase space P∂M3 for three-manifold M3 made of gluing tetrahedra. For this purpose, one first has the product of phase space of each tetrahedra ∆i ; P{∂∆i } = L Y i=1 P∂∆i (3.46) and Poisson brackets {Zi , Zj0 } = {Zi0 , Zj00 } = {Zi00 , Zj } = ~δij . (3.47) It is known that the parameter for the internal edges – CI – in (3.45) commute each other [40]. Among all edges in P{∂∆i } , the remaining linear combination of edge parameters which commute with CI parametrize the boundary phase space P∂M3 of M3 . Also, it is known that the map between e of P∂M positions and momenta (polarization {Πi }) of product phase space P∂∆i and polarization Π 3 which we choose is the affine Sp(2L, C) transformation.3 One can regard the process to obtain P∂M3 mentioned above as a symplectic quotient of P{∂∆i } with moment map CI ; P∂M = L Y i=1 ! P∂∆i // (CI = 2πi) . 2 Note that there are two edge parameters in a single tetrahedron. 3 More precisely, Sp(2L, Q) with translations by rational multiples of iπ. (3.48) 30 Regarding Lagrangian submanifold, one can do similarly. After taking products of each Lagrangian submanifolds, one eliminates algebraically the edge parameters which don’t commute with CI , then takes CI = 2πi. Then one obtains the Lagrangian submanifold LM3 of P∂M3 . One can apply above techniques similarly to 3-manifold with the geodesic surfaces boundary or the cusp boundary. If one consider a knot complement, one obtains PT 2 as a boundary phase space and one can show that from LM3 constructed via gluing procedure gives non-abelian branch of A-polynomial mentioned in section 3.2. 3.3.3 2-3 Pachner move We have discussed how boundary phase space P∂M3 and Lagrangian submanifold LM3 are obtained upon gluing tetrahedra. It is desirable that these quantities are independent of choice of triangulations of M3 . Meanwhile, it is known that any two triangulations of M3 which have same triangulations on the boundary surface of M3 are related by a sequence of 2-3 Pachner moves. 2-3 Pachner move states that two tetrahedra glued along a common face is interchangeable with three tetrahedra glued along three each faces with a common internal edge. Figure 3.3: 2-3 Pachner move Therefore, if one can show that P∂M3 and LM3 stay same under 2-3 Pachner move, then given a fixed triangulation at the boundary surface of M3 , P∂M3 and LM3 are independent with specific triangulation of M3 . This was actually shown in [29] 3.3.4 Quantization, wavefunction, and partition function One can quantize above classical analysis systematically. 4 We just quote the results here. 4 Actually, what we mean by quantization here is holomorphic quantization, so Hilbert space in this subsection is not an honest Hilbert space which is from quantization of phase space with real polarization and real symplectic form. Chern-Simons wavefunction or holomorphic block lives in the vector space of holomorphic functions which forms representation of operator algebra. But we will just use terminology “Hilbert space”. 31 b Z b0 , Zb00 , Upon quantization, edge parameters Z, Z 0 , Z 00 are promoted to quantum operator Z, b Zb0 ] = [Z b0 , Zb00 ] = [Zb00 , Z] b = ~; [Z, (3.49) and the equation for the boundary phase space becomes b+Z b0 + Z b00 = iπ + ~ Z 2 (3.50) where the ~ term is regarded as a quantum correction which is determined by requiring topological invariance of combinatorial construction of Lb and S-duality of the operator algebra. The Lagrangian submanifold becomes a quantum operator on Hilbert space, Lb∆ = zb00 + zb−1 − 1 ≃ 0 (3.51) where ≃ is understood in the sense that it annihilates a certain wavefunction. An operator algebra act on the space of locally holomorphic function in a way that Zb ≃ ~Z, ~ Zb0 ≃ iπ + − ~∂Z − Z, 2 Zb00 ≃ ∂Z , (3.52) Then up to normalization ambiguity, a wavefunction on a tetrahedron ψ(∆; Z) := ψ∆ (Z) is ψ∆ (Z) = ∞ Y r=1 (1 − q r z −1 ), (3.53) where one can check Lb∆ ψ∆ (Z) = 0. One can perform quantum gluing to obtain LbM3 from the product of a number of Lb∆ as done bI , in classical case. A bit more specifically, the internal edge parameter CI is lifted to operator C bI , and and from the product of Lb∆ ’s, one removes all elements which do not commute with the C bI = 2πi + ~. But it is more complicated due to noncommutativity. The coefficients for then takes C bI is fixed from the consideration of topological invariance of combinatorial construction of LbM . ~ in C 3 As mentioned above, in the gluing procedure, there is a map from the positions and conjugate momenta in polarization Π of the product of phase space of each tetrahedra to the positions and e of the phase space P∂M , and this map is given by the affine conjugate momenta in polarization Π 3 symplectic group ISp(2L, Q) (semiproduct of Sp(2L, Q) and affine shifts). At the quantum level, it is similar, but there is ~ correction in affine shift. We are interested in what the wavefunction will e be upon change of polarization from Π to Π. 32 We are interested in using Ψ∆ (Z) = Φ~/2 (−Z + iπ + ~/2), (3.54) for a single tetrahedron as a “wavefunction”, and how one can obtain the corresponding quantity on M3 via gluing procedure from the product of Ψ∆ (Z)’s. Here, Φ~/2 is the non-compact quantum dilogarithm [41]  Q∞ 1+qj−1/2 ep −j+1/2 p e Φ~/2 (p) = Qj=1 1+eq j−1/2 epe e  ∞ 1+eq j=1 1+q −j+1/2 ep |q| < 1 (3.55) |q| > 1 where pe = 2πi p, ~ 4π 2 qe = e− ~ . (3.56) So (3.54) can be thought as doubling of wavefunction in (3.53), which is annihilated by Lb∆ = 00 −1 b zb00 + zb−1 − 1 and Le = b ze + b ze − 1 where be b ze = exp(Z), 00 b 00 b ze = exp(Ze ), 2πi b b Z, with Ze = ~ 00 2πi b00 b Ze = Z ~ (3.57) and original variables (without tilde) and dual variables (with tilde) commute each other. Dual 00 00 variables satisfy b ze b ze = qeb zeb ze . This “wavefunction” Ψ∆ (Z) on a single tetrahedron is actually analytically continued SL(2, R) Chern-Simons partition function on a single tetrahedron. Modulo some subtleties, formally it is in Weil representation of ISp(2L, C) [42, 43, 29]. More explicitly, let ϕ be affine symplectic ISp(2L, C) action on wavefunctions ϕ Ψ(Z1 , · · · , ZL ) 7→ Ψ(E1 , · · · , C1 , · · ·) | {z } (3.58) L induced from the change of symplectic basis; ϕ∗ b ,··· ,Γ b C,1 , · · ·) b1 , · · · , Z bL , Zb100 , · · · , ZbL00 ) 7− b ,··· ,C b , · · ·, Γ (Z −→ (E |1 {z 1 } | E,1 {z } L (3.59) L bK , K = 1, · · · denote the operator for external edge (or cusp holonomy) parameters which where E e and commute with C bI , and Γ’s are corresponding conjugate are “positions” in polarization Π ⊂ Π ~b b~00 ~b ~b operators. For convenience, we call the former set as (Z, Z ) and the latter set as (X, Y ). Symplectic group Sp(2L, C) is generated by three generators and acts on the column vector 33 ~b b~00 t (Z, Z ) as a form of L by L block matrices;  I ϕ∗,T =  B 0 I  ,  I −J ϕ∗,S =  J −J I −J   , ϕ∗,U =  U 0 0 U −1 t   (3.60) where I is the identity matrix, B is symmetric matrix, J is diagonal matrices whose entries are 0 or ~ transform under above ϕ∗ as 1, and U is invertible. Then wavefunction Ψ(Z) 1 ~t ϕT ~ X ·B·X ~ 7− ~ = Ψ(X) ~ e 2~ −→ Ψ0 (X) Ψ(Z) Z ϕS ~ Z ~ 0 ~ ~ Ψ(Z) ~ e ~1 X·J ~ Ψ(Z) 7−−→ Ψ (X) ≃ dZ ϕU ~ ≃ Ψ(U −1 · X) ~ ~ 7− −→ Ψ0 (X) Ψ(Z) (3.61) (3.62) (3.63) up to overall normalization. For the affine transformation ~b b~00 ϕ∗,pos. ~b ~ (Z, Z ) 7−−−−→ (Z + ~s, Zb00 ) ~b b~00 ϕ∗,mom. ~b b~00 ~ (Z, Z ) 7−−−−−→ (Z, Z + t), (3.64) (3.65) the wavefunction f (Z) transform as ϕpos. ~ = Ψ(X ~ − ~s) Ψ(Z) 7−−−→ Ψ0 (X) ϕmom. 1~ ~ ~ = Ψ(X)e ~ ~ t·X Ψ(Z) 7−−−−→ Ψ0 (X) (3.66) (3.67) where elements in ~s or ~t are rational multiple of iπ and ~. This looks similar Sp(2L, Z) action on 3d conformal field theories with abelian symmetries discussed in Chapter 2. 34 Chapter 4 3d-3d Correspondence from Gluing Tetrahedra In similar spirit of the correspondence between 4d N = 2 superconformal field theories and Liouville (or Toda) theories on a Riemann surface [44], when 6d (2,0) theory with Lie algebra g = Lie(G) of ADE type is wrapped on three manifold − M3 − with partial twisting, we have 3d N = 2 superconformal field theories − T [M3 ; G] − described as the IR fixed points of abelian ChernSimons-matter theories determined by M3 , and if M3 has a boundary they are also determined by e of complex phase space P∂M of flat GC connection on ∂M3 ; a chosen polarization Π 3 e , G) (M3 , Π T [M3 , Π; G], (4.1) and the 3d-3d correspondence states that physical quantities such as partition functions and classical vacua of non-supersymmetric complex Chern-Simons theory on M3 are matched with those of 3d N = 2 abelian SCFT [39, 25, 45, 13, 46, 47, 48]. In this thesis, we are only interested in the case the number of brane is 2, i.e. the case GC = e G] is denoted as T [M3 ]. The case for higher rank of A-type SL(2, C), and for simplification T [M3 , Π; G was considered in [45]. Original 3d-3d correspondence in [39] was constructed from gluing ideal tetrahedra. As we discussed in previous section, one can consider Chern-Simons theory on 3-manifold obtained from gluing tetrahedra. One can also calculate classical moduli space of flat SL(2, C) vacua LM3 , and quantization LbM3 which annihilates non-perturbative Chern-Simons wavefunction or analytically continued SL(2, R) partition functions. If one can find a 3d N = 2 gauge theory corresponding to Chern-Simons theory on an ideal tetrahedron and also a proper gauge-theoretic interpretation of gluing procedure in Chern-Simons theory side, the 3d-3d correspondence is built in the construction. This is a basic idea in [39]. However, this construction doesn’t give a complete 3d-3d correspondence. This is because Chern- 35 Simons theory on three-manifold from gluing tetrahedra always missing information of abelian flat connection. Examples which captures all flat SL(2, C) connections will be discussed in Chapter 5. In this chapter, we summarize the result of the 3d-3d correspondence of [39, 25, 13]; In section 4.1, we review motivations of 3d-3d correspondence from 2d-4d correspondence and vortex partition function. In section 4.2, we summarize dictionary between Chern-Simons theory and 3d N = 2 theories. In section 4.3, we discuss known examples for 3d-3d correspondence. 4.1 Motivation for 3d-3d correspondence If M5 branes are compactified on a certain d-dimensional manifold Md with an appropriate twisting on it, one obtains 6 − d dimensional supersymmetric field theory. In this perspective, when considering a Riemann surface with some number of punctures, one obtains 4d N = 2 superconformal field theories characterized by the Riemann surface on which M5-branes are wrapped. Interestingly it was found in [44] that when the number of M5-branes is 2 the instanton partition function of 4d N = 2 SCFT is matched with conformal block of Liouville CFT (or Toda theory for N M5-branes). Such correspondence was extended to more general gauge group, also for superconformal index on S 1 ×S 3 , and with insertion of extended object such as Wilson loop, ’t Hooft loop, or surface operators. Also, the domain wall in 4d N = 4 or 2∗ theory were considered in [49, 50, 51, 11]. In geometry side, this corresponds to the mapping cylinder M3 = C ×ϕ I where the copy of a once-punctured Riemann surface C on the one side is related to the copy of a once-punctured Riemann surface C on the another side via mapping class group twist ϕ. In this setup, the duality kernel ZϕN =2 (1 , 2 ) can be interpreted as a partition function Zϕ3d (1 /2 ) of certain 3d N = 2 theory on three-sphere S 3 . In addition, the 2d-4d correspondence tells us that ZϕN =2 (1 , 2 ) = ZϕLiouville (b) (4.2) where the LHS is a partition function on the domain wall and the RHS is so called Moore-Seiberg kernel in Liouville theory with b2 = 1 /2 . In literature, it has been well-known that ZϕLiouville (b) = ZϕT eich (1 /2 ) = Z CS (Mϕ ; 1 /2 ) where ZϕT eich (1 /2 ) is a kernel in quantum Teichmüler theory associated to mapping class group ϕ and Z CS (Mϕ ; 1 /2 ) is Chern-Simons partition function on mapping cylinder Mϕ with analytically level k = 2 /1 . 36 Therefore, one can expect with above relations, Z CS (Mϕ ; 1 /2 ) = Zϕ3d (1 /2 ). (4.3) In other words, one can expect that there is a correspondence at least for Chern-Simons theory on mapping cylinder Mϕ and 3d N = 2 gauge theory on S 3 . In addition, from the “classical” limit (~ → 0) of vortex partition function of effective 2d N = (2, 2) theory [52], it was expected that (complex) Chern-Simons theory corresponds to 3d N = 2 theories in some way. For example, in loc. cit., it was discussed that moduli space of complex flat connection in Chern-Simons theory on 3-manifold is equal to the moduli space of supersymmetric vacua of effective 2d N = (2, 2) theories. So these make us to expect that there is a correspondence between Chern-Simons theory on 3-manifold and 3d N = 2 theories. 4.2 3d-3d correspondence of Dimofte-Gaiotto-Gukov In this section, we summarize the dictionary between 3d N = 2 theories and Chern-Simons theory on M3 admitting tetrahedra triangulation [39, 25, 13]. 4.2.1 3d-3d correspondence for a tetrahedron We first consider a 3d N = 2 theory for a single tetrahedra. Regarding partition functions, we saw that sine double function appeared in partition function on Sb3 of 3d N = 2 theories. At the same time, as mentioned in section 3, the non-compact quantum dilogarithm function, which is closely related to sine double function, appeared in a partition function of SL(2, R) Chern-Simons theory. From this observation, one can read off field contents and Chern-Simons coupling corresponding to a single tetrahedron by matching variables between two such as a certain quantity of a chiral multiplet and an edge parameter. At the same time, one can also check with moduli space of suerpsymmetric vacua (or supersymmetric parameter space) of 3d N = 2 theories and the moduli space of flat SL(2, C) connection of Chern-Simons theory on a single tetrahedron. In sum, the 3d-3d correspondence for a single tetrahedron M3 = ∆, • A 3d N = 2 theory T [∆; Π] associated to a single tetrahedron ∆ and polarization Π of P∂∆ is a free chiral multiplet charged +1 under background U (1) gauge symmetry (i.e. flavor or global symmetry) with Chern-simons level −1/2. Given a tetrahedron ∆, if choosing a polarization as ΠZ = (Z, Z 00 ), a free chiral Φ with U (1)Z corresponds to an edge parameter Z of ∆, which we denote such chiral as ΦZ . Then Re(Z) and 37 Im(Z)/π corresponds to the twisted mass parameter of U (1)Z background gauge symmetry or global symmetry and R-charge of ΦZ , respectively. • Lagrangian submanifold L∆ in classical phase space P∂∆ , which is classical moduli space of flat SL(2, C) connections, corresponds to supersymmetric parameter space of T [∆] on R2 × S 1 ; Mflat (∆; SL(2, C)) = {z 00 + z −1 − 1 = 0} = MSUSY (∆) (4.4) • Partition function on squashed 3-sphere Sb3 , superconformal index on S 2 ×q S 1 , and holomorphic block D2 ×q S 1 of T [∆] are equal to partition function of SL(2, R) Chern-Simons theory, partition function of SL(2, C) Chern-Simons theory, and analytically continued Chern-Simons wavefunction on ∆, respectively; ZSb3 (T [∆]) = ZS 2 ×q S 1 (T [∆]) = ZD2 ×q S 1 (T [∆]) = 1−q r+1 z −1 r=0 1−(Lq)r (Lz)−1 Q∞ 1−qr+1 z−1 r=0 1−q r z Q∞ r+1 −1 z ) r=0 (1 − q Q∞ = ZCS (∆; SL(2, R)) (4.5) = ZCS (∆; SL(2, C)) = ψCS (∆) L where q = e~ . For the Sb3 partition function, ~ = 2πib, L~ = 2πib−1 with Lq = e ~ , and b is a squash parameter of Sb3 ; Lz = z 2πi/~ with z = eZ and Z = 2πbmZ + iπ + ~2 RZ where mZ and RZ are real mass and R-charge of ΦZ . For the index, z = q m/2 ζ, z = q m/2 ζ −1 with m and ζ being magnetic flux and fugacity (also called as chemical potential) for U (1)Z symmetry. In addition, quantum operator Lb∆ discussed in section 3.3 is interpreted as Ward identity for line operators in a 4d theory coupled to the boundary theory T [∆]. A bit more explicitly, if M3 has geodesic boundary C = ∂M3 , T [M3 ] provide a half-BPS boundary condition of N = 2 abelian theory on Coulomb branch of 4d N = 2 T [C, su(2)] theory, i.e. Seiberg-Witten theory of T [C, su(2)]. Then line operators in 4d N = 2 abelian theory, which are brought into the 3d boundary where T [M3 ] live, satisfy Ward identity. For example, when M3 = ∆ with polarization ΠZ , z 00 + z −1 − 1 ≃ 0 is interpreted as Ward identity H + W −1 − 1 ≃ 0 of Wilson (W) and ’t Hooft (H) line operator satisfy where z, z 0 , and z 00 correspond to Wilson, Wilson-’t Hooft, and ’t Hooft line operator, respectively [39]. We have seen how 3d-3d correspondence works for a single tetrahedra. In order to have the correspondence for general 3-manifold M3 obtained by gluing tetrahedra, we also need to match gluing procedures of both sides. This is successful [39, 25, 13] as we will review in next subsection, so 3d-3d correspondence holds for general 3-manifold obtained by gluing tetrahedra. 38 4.2.2 Gluing procedures Given a correspondence above, one can expect that (affine) SL(2, Z) action on 3d N = 2 superconformal field theories with an abelian symmetry correspond to (affine) SL(2, Z) action on the polarization Π of P∂∆ . More generally, (affine) Sp(2L, Z) action on 3d N = 2 superconformal field theories with U (1)L symmetries is expected to correspond to (affine) Sp(2L, Z) action which maps e of P∂M . It turns out that one can match the polarization Π1 × · · · × ΠL of P{∂∆i } to polarization Π 3 each other. More specifically, gluing procedure for 3d N = 2 theories T [∆i ; Π], i = 1, · · · L for L tetrahedra is constructed as follows; 1. One first takes a product of theories T [∆i , Πi ]; T [{∆i }, {Πi }] = T [∆1 , Π1 ] ⊗ · · · ⊗ T [∆L , ΠL ]. The resulting theory T [{∆i }, {Πi }] consists of L chiral multiplets Φi corresponding to edge parameter Zi of product polarization Π1 × · · · ΠL of product boundary phase space P∂∆1 × · · · × P∂∆L , and each of which are charged +1 under U (1)L := U (1) × · · · × U (1) global symmetries. There are | {z } also Chern-Simons term with level − 21 for each U (1)L . L e for the boundary phase space P∂M of M3 such that positions 2. One chooses a new polarization Π 3 e and in a new polarization and momenta in Π = Π1 ×· · ·×ΠL maps to positions and momenta in Π e Quantization of positions the internal edge CI is given by linear combinations of positions of Π. e are related to those in Π1 × · · · × ΠL by affine symplectic transformation and momenta in Π h 2L i b1 , · · · , ZbL , Z b00 , · · · Z b00 )T of Π1 × also acting on 2L column vector (Z Sp(2L, Z) n (iπ + ~2 )Z 1 L · · · × ΠL . 3. One applies the affine symplectic action g ∈ Sp(2L, Z) n h 2L i (iπ + ~2 )Z on product theories T [∆i ], which also appeared in Chern-Simons theory as Weil transformation   U 0  with invertible U ∈ GL(L, Z) act on U (1)L vector 3a. GL-type action gU =  0 U −1 t multiplets by linear transformation so that it refines U (1)L global symmetries.   I 0  with symmetric matrix B add (mixed) Chern-Simons term for 3b. T -type action gT =  B I background vector multiplet with level kij = Bij .   I −J −J  with a diagonal matrix J = diag(j1 , · · · , jL ) with ji 3c. S-type action gS =  J I −J is 0 or 1 gauge background U (1)i symmetry and introduce topological U (1) symmetry with respect to U (1)i for i such that ji = 1. 3d. Affine shifts are relevant for the theories on compactified spaces such as D2 ×q S 1 , Sb3 , or S 2 ×q S 1 , but not on R3 . Affine shifts on the positions add a unit of flavor symmetry to 39 R-symmetry. In terms of holomorphic block, it shifts Wilson line for flavor symmetry by 1 −iπ − ~2 = log(−q 2 ). Affine shifts on the momenta add mixed Chern-Simons term between the background U (1)R symmetry and global symmetry. 4. We add a superpotential W = P I OI where each operator OCI is a product of chiral fields, which comes from the internal edge parameter CI given by linear combination of edge parameters for 1 positions. This breaks all U (1) global symmetries associated to CI for each I, and U (1) 2 dimP∂M are left. Thus from above gluing procedure, one can construct 3d N = 2 theories corresponding to ChernSimons theory on M3 admitting tetrahedra triangulations. Also, as the quantities are matched in the case of a tetrahedron and gluing procedure is identified, one can calculate a supersymmetric parameter space, Ward identity of line operators, partition function on Sb3 , index on S 2 ×q S 1 , and holomorphic block of T [M3 ] via gluing procedure as done in section 3.3 for Chern-Simons theory1 and by construction they correspond, respectively, to classical moduli space of flat SL(2, C) connections LM3 , quantum operator LbM3 obtained by quantization of LM3 , SL(2, R) partition function ZCS (M3 ; SL(2, R)), full SL(2, C) partition function ZCS (M3 ; SL(2, C)), and analytically continued Chern-Simons wavefunction ψCS (M3 ). 4.3 Examples We would like to provide briefly some known examples from [39, 25, 13]. 4.3.1 SQED with Nf = 1, XYZ model, and 3d N = 2 mirror symmetry The SQED with one flavor (Nf = 1) in 3d N = 2 theory is U (1) gauge theory with two chiral mulitplets charged oppositely under gauge group. They have same charge under axial U (1) global symmetry. For the XYZ model, there is no gauge symmetry, and there are three chiral multiplets with superpotential given by product of three chirals. Firstly, we would like to see how they arise from gluing tetrahedra. 4.3.1.1 SQED Nf = 1 and XYZ model from gluing tetrahedra XYZ model is made from gluing three tetrahedra with two adjacent faces of each three tetrahedra are glued in a way that they have a common internal edge. One choose a polarization 1 The full SL(2, C) Chern-Simons partition function on M via gluing procedure was not discussed in section 3.3, 3 but it can be done similarly and was discussed in detail in section 6.2 of [25]. 40 (b) SQED Nf = 1: gluing 2 tetrahedra (a) XYZ model: gluing 3 tetrahedra Figure 4.1: Gluing tetrahedra e = (X1 , X2 , C; P1 , P2 , Γ) which are related to position-momentum (Z, Z 00 ), (W, W 00 ), and (Y, Y 00 ) Π of Π1 × Π2 × Π3 by   1       X2  0        C  1  =     P1  0        P2  0    Γ 0 X1  0 0 0 0 1 0 0 0 1 1 0 0 0 0 1 0 0 0 0 1 0 0 0 0  1   0  0   0  0   −1 0   −1 0  1 0 0 0 0 0 0 1 0 0 0 0 1 0 0 0 0 1 0 0 0 0 1 0 1 0 0 0   Z  0           0  W   0          0  Y   0    +     00   0 Z   iπ + ~2          0  W   iπ + ~2      −iπ − ~2 1 Y 00 (4.6) where the matrix on the right is GL-type and the one on the left is T -type. From the product of three Lagrangian (in flat spacetime) 1 L{Πi } [VZ , VW , VY ] = 4π Z 1 1 1 d θ − ΣZ VZ − ΣW VW − ΣY VY 2 2 2 4 Z + d4 θ Φ†Z eVZ ΦZ + Φ†W eVW ΦW + Φ†Y eVY ΦY (4.7) By applying the dictionary, adding Chern-Simons term for VY (T -type) and redefining VZ = VX1 , VW = VX2 , and VY = VC − VX1 − VX2 , one obtains LΠ e [VX1 , VX2 , VC ] = L{Πi } [VX1 , VX2 , VC − VX1 − VX2 ] + 1 4π Z d4 θ(ΣC − ΣX1 − ΣX2 )(VC − VX1 − VX2 ) (4.8) From VY = VC − VX1 − VX2 , there is a superpotential W = µΦX ΦY ΦZ . One can do similarly for SQED with Nf = 1 which is obtained from gluing two tetrahedra glued along one face, but we don’t discuss it further. Summarizing the field contents and Chern-Simons levels, with relabeling parameters for later use 41 [13], theory TXY Z of XYZ model is charges : Φ1 Φ2 Φ3 U (1)x 1 0 −1 U (1)y 0 1 −1 U (1)R 0 0 2 CS levels : U (1)x U (1)y U (1)R U (1)x 0 1/2 0 U (1)y 1/2 0 0 U (1)R 0 0 −1/2 (4.9) with a superpotential W = µ Φ1 Φ2 Φ 3 . (4.10) U (1)x and U (1)y are flavor symmetries whose complexified twisted mass parameters are X = log x and Y = log y, respectively. Above Chern-Simons level above are UV Chern-Simons level. For SQED with Nf = 1, charges : ϕ1 ϕ2 v+ v− U (1)s U (1)x U (1)y U (1)R U (1)s 1 −1 0 0 U (1)s 0 1/2 1 −1 U (1)x 0 1 0 −1 U (1)x 1/2 −1/2 0 1/2 U (1)y 0 0 1 −1 U (1)y 1 0 0 0 U (1)R 0 0 0 2 U (1)R −1 1/2 0 −1 (4.11) CS levels : without superpotential. U (1)s denotes gauge symmetry, and U (1)x is flavor symmetry and U (1)y is a topological symmetry where this notation with label x and y is chosen because they are related to those of XYZ model via 3d N = 2 mirror symmetry. Here, v+ and v− are gauge invariant monopole operators. 4.3.1.2 Holomorphic block, partition function on Sb3 , and index R For general R-charge of chiral multiplet, a tetrahedron theory T∆ charged under U (1)y global flavor symmetry has the following charges and Chern-Simons level; Φ charges : U (1)y U (1)R (1 − R)/2 U (1)y +1 U (1)y −1/2 U (1)R R U (1)R (1 − R)/2 −(1 − R)2 /2 (4.12) 42 R Holomorphic block of T∆ is given by (R) B∆ (y; q) = ((−q 1/2 )2−R y −1 ; q)∞ (4.13) Chern-Simons term of U (1)y gauge multiplet with level −1 are encoded in theta function θ(x; q) = (−q 1/2 x; q)∞ (−q 1/2 x−1 ; q)∞ . (4.14) More generally, if we have theory with Chern-Simons level matrix kij , i, j = 1, · · · p of U (1)p gauge/global symmetries and mixed Chern-Simons level σi , i = 1, · · · p with U (1)p symmetries and U (1)R symmetry, then mixed CS levels kij and σi are encoded in product of theta functions; nh Y θ (−q 1/2 )bh ~y ~ah ; q (4.15) h where bh and nh are integers for each h, ~y = (y1 , · · · , yp ), and ~ah = (ah,1 , · · · , ah,p ) are vectors of p integers for each h. When performing fusion, for S-fusion 1 ~ 2 θ((−q 1/2 )b y a ; q) S = i] C [ exp − (a · X)2 + (iπ + ) b (a · X) 2~ 2 (4.16) where x = exp X, x e = exp 2πi ~ X and ] and [ are certain number. For id-fusion, 2 θ((−q 1/2 )b y a ; q) id = (−q 1/2 )−(a·m)b ζ −(a·m)a (4.17) where x = q m/2 ζ and x e = q m/2 ζ −1 . From this information, holomorphic block of TXY Z is given by BXY Z (x, y; q) = (qx−1 ; q)∞ (qy −1 ; q)∞ (xy; q)∞ θ(−q −1/2 xy; q) (4.18) Holomorphic block of SQED with Nf = 1 is ds ΥSQED (s, x, y; q) 2πis Γ Z ds θ((−q −1/2 )y; q) (qs−1 ; q)∞ (qx−1 s; q)∞ = −1/2 )sy; q) Γ 2πis θ((−q Z BSQED (x, y; q) = (4.19) (4.20) where Γ is an integration cycle we would like to find. There is only one critical point of twisted 43 superpotential ~ log ΥSQED (s, x, y; q), which is s(1) = (y − x−1 )/(y − 1). Appropriate integration cycle Γ depends on the sign of ~ and value of x and y. We don’t cover all possibilities here, but for example when ~ > 0 (i.e. |q| > 1) and for certain value of x and y, we can choose a contour enclosing poles from (qs−1 ; q)∞ or (qx−1 s; q)∞ . 2 For either |~| > 0 or |~| < 0, it can be shown that exactly or numerically [13] that two holomorphic blocks from XYZ model and SQED with Nf = 1 are same up to overall q-dependent factors. Modulo q-dependent factor for S-fusion, partition function on Sb3 or index agree; Z 2 BXY Z 2 = (4.21) dσ 2 2 ΥSQED id = BSQED id S 1 2πiσ (4.22) R BXY Z 2 2 dS ΥSQED S = BSQED S Z = where S = log s for S-fusion and s = q m/2 ζ for id-fusion. This checks that XYZ model and SQED with Nf = 1 are 3d N = 2 mirror dual. 4.3.1.3 Supersymmetric parameter space fSQED /∂σ = 0 for σ, putting the solution σ = σ (1) back to W fSQED , and by solving After solving ∂ W fSQED (σ (1) )/∂X1 , P2 = ∂ W fSQED (σ (1) )/∂X2 , we obtain the supersymmetric parameter P1 = ∂ W space p1 + p2 − 1 = 0, x1 p2 + p1 −1 x2 =0 (4.23) fXY Z , one obtains Meanwhile, for XYZ model with the twisted superpotential W γp1 + 1 − 1 = 0, x1 γp2 + 1 − 1 = 0, x2 − x1 x2 γx1 x2 + −1=0 c c (4.24) fXY Z /∂X1 , P2 = ∂ W fXY Z /∂X2 , and Pγ = ∂ W fXY Z /∂Γ where upper case letter is from P1 = ∂ W log of lower case letter and P1 , P2 , and Γ are interpreted as effective FI parameters with respect to U (1)X1 , U (1)X2 , and U (1)C . The condition C = 2πi gives c = 1. Eliminating γ, we obtain the moduli space of supersymmetric vacua p2 − 1 = 0, (x1 − 1) p1 + x1 p1 (x2 − 1) p2 + − 1 = 0. x2 (4.25) Here, x1 = 1 or x2 = 1 i.e. X1 = 0 or X2 = 0 corresponds to singular locus since they mean that the mass of chiral multiplets are zero, which we should not integrate out. So for x1 , x2 6= 1, they are equivalent. 2 Note that these expression is valid for |q| < 1. For |q| > 1 one has 1/(s−1 ; q −1 ) −1 s; q −1 ) , respectively. ∞ and 1/(x ∞ 44 4.3.2 Trefoil knot and figure-eight knot In this subsection, we quote results from trefoil knot (31 ) and figure-eight knot (41 ) complement in S 3 [39, 25, 13]. One can glue two tetrahedra to make trefoil knot complement in S 3 . By similar way above, one obtains T DGG [31 ] T DGG [31 ] : Φ1 Φ2 U (1)s 0 0 U (1)x 1 −1 U (1)R 2 0 , CS : U (1)x U (1)R U (1)x 3 3 U (1)R 3 ∗ . (4.26) Thus, holomorphic block of T DGG [31 ] is simple and is given by θ(x; q)−3 . So partition functions on Sb3 and index are readily calculated. The moduli space of supersymmetric vacua is y + x3 = 0 (4.27) which is A-polynomial of trefoil knot for irreducible flat SL(2, C) connection.3 Actually, above T DGG [31 ] is simplified theory, because we don’t have enough superpotential to break a flavor symmetry, which is turned off by hand in above gluing ([25, Section 4.3] for detail). But since partition functions on Sb3 , index, and moduli space of supersymmetric vacua are insensitive to superpotential, above theory is good enough for calculating them. For figure-eight complement, T DGG [41 ] : Φ1 Φ2 U (1)s 1 1 U (1)x 1 −1 U (1)R 0 0 . (4.29) and there is no UV Chern-Simons term. This is also a simplified theory, but there is refined triangulations with 6 tetrahedra, which have enough superperpotential to break flavor symmetries ([39, Section 4.6] for detail). One can calculate holomorphic block, partition function on Sb3 , and superconformal index, which 3 Notation l and m for longitude and meridian eigenvalues are related to y and x as l ↔ y, m2 ↔ x (4.28) 45 we don’t calculate here. The moduli space of supersymmetric vacua is y 2 − (x2 − x − 2 − x−1 − x−2 )y + 1 = 0, which is A-polynomial of 41 for non-abelian flat SL(2, C) connection. (4.30) 46 Chapter 5 Toward a Complete 3d-3d Correspondence 5.1 3d-3d correspondence revisited So far we have summarized each side of the correspondence and discussed dictionary between ChernSimons theory and 3d N = 2 theory. However, as we have mentioned, Chern-Simons theories on 3-manifold obtained from gluing ideal hyperbolic tetrahedra do not capture abelian branch. Thus, the corresponding 3d N = 2 theories corresponding to gluing tetrahedra do not capture such branch. This can be seen in supersymmetric parameter space, partition functions, or superconformal index. 5.1.1 M-theory perspective To better understand the problem, it is useful to consider M5-brane systems giving 3d-3d correspondence. When N M5 branes are wrapped on 3-manifold M3 , we obtain the theory T [M3 ] on R3 part (or S 2 × R, D2 × R) of M5 branes worldvolume in the following M5-brane system; space-time: R5 × ∪ N M5-branes: R3 CY3 ∪ × (5.1) M3 where M3 is embedded in a Calabi-Yau 3-fold CY3 as a special Lagrangian submanifold. In order to make T [M3 ] be supersymmetric, there should be appropriate topological twisting on M3 . There are usually two choices on Calabi-yau 3-fold. One is a cotangent bundle of M3 , T ∗ M3 , and another is resolved conifold X which is O(−1) ⊕ O(−1) bundle over CP1 . The latter is relevant to physical realization of knot homologies. In the study of knot homologies for knot complement in 47 S 3 , M5-brane system is space-time: R5 × X ∪ M5-brane: R3 ∪ × (5.2) eK L e K is a Lagrangian submanifold in resolved conifold X. Here, L In the standard 3d-3d correspondence, we are more interested in the case where Calabi-Yau 3fold is a cotangent bundle T ∗ M3 of M3 in (5.1). More specifically, in this thesis we are interested in a knot complement in S 3 − S 3 \K − and the number of M5-branes is two. Therefore we take M3 in as 3-sphere S 3 , so we consider CY3 = T ∗ S 3 . In addition, in order to incorporate knot K, we also introduce a single M5-brane. The resulting M5-brane system is space-time: R5 × ∪ N M5-branes: R3 M5-brane: R3 T ∗S3 ∪ × S3 × LK (5.3) where LK is a conormal bundle to the knot K [53], which is a lagrangian submanifold in T ∗ S 3 such that LK ∩ S 3 = K (5.4) Here, a single M5-brane corresponds to knot in S 3 , and we can regard that LK creates a monodromy defect along K in S 3 . So in this setup 3d-3d correspondence is about SL(2, C) Chern-Simons theory on a knot complement in S 3 − S 3 \K − and 3d N = 2 theory T [S 3 \K]. For partition function on squashed 3-sphere Sb3 , superconformal index on S 2 ×q S 1 , and holomorphic block on R2q × S 1 of T [S 3 \K], R3 in above M5-brane system can be taken as Sb3 , S 2 ×q S 1 , or R2q × S 1 . As a side remark, M5-brane system for knot homologies in (5.2) can be thought as large N dual of M5-brane system in (5.3); for example, under geometric transition, Lagrangian submanifold LK e K in a resolve conifold X [54]. in T ∗ S 3 maps to Lagrangian submanifold L 5.1.1.1 Symmetries of T [M3 ] from the perspective of M5-brane system Back to the most general M-theory system, we would like to symmetries in (5.1). There are at least three U (1) symmetries which are independent of the specific choice of M3 . One is a Cartan subgroup of the SO(3) rotation symmetry of R3 . Another is a rotation symmetry in two-dimensional transverse space of R3 in R5 . We also have U (1) R-symmetry. Certain linear combinations of these 48 three U (1) symmetries give three conserved charges of the gauge theories. This is familiar to the case of surface operator in 4d N = 2 gauge theory where there are three fugacities in the superconformal index in the presence of surface operator whose nature is independent of specific choice of surface operators and the theory that surface operator is present. These three U (1) symmetries should appear in 3d theory T [M3 ] as T [M3 ] should exhibit all properties the M5-brane system. Firstly, U (1) rotational symmetry on R3 is obviously a part of Lorentz symmetry. And certain linear combination of other two U (1) symmetries become U (1)R symmetry of T [M3 ]. Another linear combination gives non-R global symmetry which we called as U (1)t . This U (1)t symmetry has not appeared in the previous 3d-3d correspondence, or usual 3d N = 2 theories. However, since this is a symmetry of M5-brane system (5.1), theories T [M3 ] should have it as well. Actually, this U (1)t symmetry plays a key role in the 3d-3d correspondence as we will see soon. 5.1.1.2 Supersymmetric parameter space As mentioned in section 4.1, in [52], by compactifying M5-brane system on M-theory circle of R2q ×S 1 (instead of “R3 ” and similarly for R5 in (5.1)), it was shown that with appropriate twist on M3 the supersymmetric parameter space of effective 3d N = 2 theory which is regarded as 2d N = (2, 2) theory on R2q is the moduli space of GC flat connection on M3 ; MSUSY (T [M3 ; G]) = Mflat (M3 ; GC ) (5.5) What we have seen above for the previous 3d-3d correspondence by Dimofte-Gaiotto-Gukov was that MSUSY (TDGG [M3 ; G]) 6= Mflat (M3 ; GC ) (5.6) That is because abelian flat GC connection on M3 is captured via gluing tetrahedra. So, more precisely, DGG MDGG flat (M3 ; GC ) 6= Mflat (M3 ; GC ) = MSUSY (T [M3 ; G]) 6= MSUSY (T [M3 ; G]) (5.7) DGG where MDGG flat (M3 ; GC ) = MSUSY (T [M3 ; G]). What we would like to do is to find several simple examples such that (5.5) actually hold for G = SU (2). 5.1.2 Brief comments on recent developments in 3d-3d correspondence There also have been some number of developments in 3d-3d correspondence, and they all imply that the gauge theory T [M3 ] should realize all GC flat connections on M3 . They include 49 • The recent proof [47, 48] of the 3d-3d correspondence indicates that all complex flat connections on M3 should be treated democratically and, therefore, no one should be left behind. • Various deformations / refinements of (5.5) necessarily require taking a proper account of all branches [37, 55, 56, 57], and can serve as a useful tool in identifying T [M3 ] even in the undeformed case. This will be our approach in this thesis. • The correspondence [58] between 4-manifolds and 2d N = (0, 2) theories T [M4 ; G] represents gluing 4-manifolds along M3 as a sequence of domain walls and boundary conditions in 3d N = 2 theory T [M3 ; G]. Much like the other developments, it works only if all GC flat connections on M3 are realized in T [M3 ; G]. 5.2 Goal and strategy 5.2.1 T [M3 ] from homological knot invariants and Higssing We would like to construct some theories T [M3 ; G] with all expected flavor symmetries and with vacua corresponding to all flat connections on M3 , and to investigate their relation to theories TDGG [M3 ; G] . We will mainly focus on the case G = SU (2), and on knot complements M3 = S 3 \K. A knot-complement theory T [M3 ] := T [M3 ; SU (2)] is defined by compactification of the 6d (2,0) theory on S 3 with a codimension-two defect wrapping the knot K ⊂ S 3 . In this case T [M3 ] should gain a U (1)x flavor symmetry, part of the SU (2)x flavor symmetry of the defect, in addition to U (1)t and U (1)R . What we find can be then summarized by the following diagram: T [M3 ] h∂r Ox i6=0 . U (1)x Tpoly [M3 ; r]——- &hOt i6=0 (5.8) U (1)t TDGG [M3 ]——- In particular, the theory TDGG [M3 ] is a particular subsector of T [M3 ] obtained by Higgsing the U (1)t symmetry. The left-hand side of the diagram (5.8) indicates an expected relation between T [M3 ] and a theory Tpoly [M3 ; r] whose partition functions compute the Poincaré polynomials of r-colored SU (2) knot homology for K. Indeed, our practical approach to constructing T [M3 ] will be to identify a 3d N = 2 theory with U (1)x ×U (1)t symmetry whose partition functions reduce to the desired Poincaré polynomials in a special limit. Physically this limit corresponds to another Higgsing procedure, this time breaking the U (1)x symmetry of T [M3 ] while creating a line defect or vortex, similar to scenarios studied in [59, 60, 61]. An important feature of (5.8) is that the two arrows corresponding to Higgsing do not commute. In particular, while it is easy to obtain Jones polynomials of knots from the Poincaré polynomials 50 on the left-hand side by ignoring U (1)t fugacities, it is (seemingly) impossible to do this from TDGG [M3 ] on the right-hand side. Jones polynomials include a crucial contribution from the abelian flat connection on a knot complement M3 as discussed in section 3, and vacua corresponding to the abelian flat connection are lost during the Higgsing of U (1)t . 5.2.2 Boundary of M3 and T [M3 ] Later, in section 5.6 we discuss gluing of knot and link-complement theories to form closed M3 , in particular 3d N = 2 theories for lens spaces, and Brieskorn spheres. The importance of such gluing or surgery operations is two-fold. First, it will give us another clear illustration why all flat connections need to be accounted by 3d N = 2 theories T [M3 ] in order for cutting and gluing operations to work. Moreover, it will help us to understand half-BPS boundary conditions that one needs to choose in order to compute the half-index of T [M3 ]. As explained in [58], a large class (“class H”) of boundary conditions can be associated to 4-manifolds bounded by M3 , 4-manifold M4 bounded by M3 boundary condition for 3d N = 2 theory T [M3 ] (5.9) 2d wall 3d T− 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 1111 0000 a) 3d T+ T− 2d wall therefore making the half-index of T [M3 ] naturally labeled by 4-manifolds. T+ b) Figure 5.1: The index of 3d N = 2 theories can be generalized to include domain walls and boundary conditions [62]. It is obtained from two copies of the half-index IS 1 ×q D± (T ± ) ≃ Zvortex (T ± ) 1 convoluted via the index (flavored elliptic genus) of the wall supported on S 1 × Seq , where D± is 2 1 + − the disk covering right (resp. left) hemisphere of the S and Seq := ∂D = −∂D is the equator of the S 2 . 51 5.3 Contour integrals for Poincaré polynomials Even though our main goal is to identify all symmetries and flat connections in the 3d N = 2 theory T [M3 ], one of the intermediate steps is of mathematical value on its own. Namely, the point of this section will be to show that Poincaré polynomials of homological link invariants can be expressed as contour integrals ds Υ(s, q, t, . . .) Γ 2πis Z PK (q, t . . .) = (5.10) in complex space C∗ m parametrized by (multi-)variable s. Here, PK (q, t . . .) stands for the Poincaré polynomial of a doubly-graded [63, 64, 65, 66] or triply-graded [67, 68, 69] homology theory H(K) of a link K: PK (q, t . . .) = X i,j,... q i tj . . . dim Hi,j (K) (5.11) that categorifies either quantum sl(N ) invariant [31] or colored HOMFLY polynomial [70], respectively. Depending on the context and the homology theory in question, the sum runs over all available gradings, among which two universal ones — manifest in (5.11) — are the homological grading and the so-called q-grading. In the case of HOMFLY homology, there is at least one extra grading and, correspondingly, the Poincaré polynomial depends on one extra variable a, whose specialization to a = q N makes contact with sl(N ) invariants. The Poincaré polynomials of triply-graded HOMFLY homology theories are often called superpolynomials. In general, such invariants are also labeled by a representation / Young diagram R and referred to as colored, unless R = in which case the adjective ‘colored’ is often omitted. In this section we will write the Poincaré polynomials of colored knot homologies in the form (5.10) of contour integrals, whose physical interpretation will be discussed in the later sections. Our basic examples here and the rest part of this thesis will be the unknot, trefoil, and figure-eight knot complements. In general, superpolynomials or Poincaré polynomials are expressed as finite sums of products of q-Pochhammer symbols (z; q)n := n−1 Y i=0 (1 − q i z) (5.12) and monomials. For instance, the unnormalized superpolynomial1 of the trefoil 31 is [38] (see also [69, 71, 72, 73]): Sr P 31 (a, q, t) = r X (a(−t)3 ; q)r (−aq −1 t; q)k k=0 (q; q)k (q; q)r−k r r 3r a 2 q − 2 q (r+1)k (−t)2k− 2 . (5.13) 1 We can proceed with normalized superpolynomial, which is obtained from dividing unnormalized superpolynomial by superpolynomial of unknot. However, since physical interpretation of unnormalized superpolynomial is clearer than than normalized one, so we focus on unnormalized superpolynomial. 52 This is the Poincaré polynomial of the HOMFLY homology (5.11) colored by the r-th symmetric power of the fundamental representation of SU (N ) or, in the language of Young diagrams, by a Young tableau with a single row and r boxes. For our applications here, we specialize to SU (2) homology2 by setting a = q 2 . It is further convenient to renormalize the SU (2) polynomial by a factor (−1)r , defining3 Sr P3r1 (t; q) := (−1)r P 31 (a = q 2 , q, t) = r X (q 2 (−t)3 ; q)r (q(−t); q)k (q; q)k (q; q)r−k k=0 1 (−q 2 )2rk+2k+r (−t)2k−2r . (5.14) We remark that the following steps could also be carried out for generic a, though for our applications we specialize from SU (N ) to SU (2). Let us suppose that |q| > 1 (for reasons that will become clear momentarily), and define −1 (z)− )∞ = ∞ := (z; q ∞ Y (1 − q −i z) , 1 1 − 2 −1 − θ− (z) := θ(z; q −1 ) = (−q − 2 z)− z )∞ , ∞ (−q (5.15) i=0 as well as θ− (z1 , ..., zn ) := θ− (z1 ) · · · θ− (zn ) . r − Then, by using identities such as (q r z)− ∞ /(z)∞ = (qz; q)r = (−1) q θ− (q n z)/θ− (z) = q n2 2 =: r(r+1) 2 z r (q −1 z −1 ; q −1 )r and z n , we may rewrite ∞ P3r1 (t; q) = (5.16) 3 1 3 3 − X (−1/(q 2 t3 ))− (s/(qx))− θ− (q 2 sxt3 , −q 2 x, −q 2 x(−t) 2 , 1) ∞ (−1/(qt))∞ ∞ 3 1 3 3 − 2 3 − (q −1 )− ∞ (−1/(q xt ))∞ k=0 (q; q)k (−1/(qst))∞ θ − (q 2 xt3 , −q 2 x/s, −q 2 (−t) 2 , x) x=q r, s=q k ∞ X 1 (0) Υ31 (s, x, t; q) x=qr, s=qk . −1 )− (q; q) (q ∞ k k=0 (5.17) Note in particular that upon setting x = q r and s = q k the term (s/(qx))− ∞ in the numerator on the LHS vanishes unless k ≤ r. Thus the sum naturally truncates to the one in (5.14). Going further, we observe that the sum in (5.17) may be rewritten as a sum of residues P3r1 (t; q) = X ∞ k=0 Ress=qk 1 1 (0) Υ (s, x, t; q) , 31 2πis (s)− ∞ x=q r (5.18) (0) k −1 − since Υ31 is smooth at s = q k , while the residue of 1/[2πis(s)− )∞ ]. ∞ ] at s = q is precisely 1/[(q; q)k (q It was the initial choice |q| > 1 that allowed us to write the sum as residues like this. Therefore, at 2 Specialization a = q 2 leads to Poincaré polynomials of colored SU (2) knot homologies for a certain class of knots, which include unknot, trefoil, and figure-8 knot considered in this thesis. In general and for more complicated knots, specialization of superpolynomials to Poincaré polynomials of SU (N ) knot homologies requires taking into account a nontrivial action of differentials [67, 69, 37]. 3 In the next section, the rescaling by (−1)r leads to a convenient choice of fermion-number twist when identifying P3r1 (t; q) with a partition function of T [31 ] on R2 ×q S 1 . 53 least formally, P3r1 (t; q) = ds Υ31 (s, x, t; q) 2πis ΓI x=q r Z (5.19) with Υ31 (s, x, t; q) := 1 (0) Υ31 (s, x, t; q) (s)− ∞ 1 3 (5.20) 3 3 − θ− (q 2 sxt3 ) θ− (−q 2 x, −q 2 x(−t) 2 , 1) (−1/(q 2 t3 ))− ∞ (−1/(qt))∞ × = 3 3 3 − − − , (−1/(q 2 xt3 ))∞ (s)∞ (−1/(qst))− θ− (q 2 xt3 , −q 2 (−t) 2 , x) ∞ (x/s)∞ where the contour ΓI is shown in Figure 5.2. (We have put all s-dependent terms in Υ31 on the right.) This is now the form of a contour integral (5.10). 1/(xt3 ) ΓII 1/(qt) ΓII 1/(xt3 ) ΓI x → qr 1 x ΓI 1 qr ΓIII arg s ΓIII 1/(qt) log |s| Figure 5.2: Possible integration contours for the trefoil, drawn on the cylinder parametrized by log s. There are three half-lines of poles in the integrand Υ31 (s, x, t; q), coming from 3 − − − 32 (s)− ∞ , (−1/(qst))∞ , (x/s)∞ in the denominator; and a full line of zeroes from θ (q sxt ) in the numerator. On the right, we demonstrate a pinching of contours as x → q r . − − Note that the three terms (s)− ∞ , (−1/(qst))∞ , and (x/s)∞ each contribute a half-line of poles to Υ31 . If we take q > 1 to be real, then the asymptotics of the integrand are given by   exp 1 log q (log x + 3 log(−t)) log s + . . . log |s| → ∞ Υ31 ∼  exp 1 (− 1 (log s)2 + . . . log |s| → −∞ , log q 2 (5.21) so the integral along ΓI does converge in a suitable range of x and t (namely, if |xt3 | < 1). In contrast, the integrals along the other obvious cycles here, ΓII and ΓIII , always converge. Moreover, a little thought shows that upon setting x = q r the integral along ΓI must equal the integral along ΓIII ; indeed, as x → q r , some r + 1 pairs of poles in the lines surrounded by ΓI and ΓIII collide, and all contributions to the integrals along either ΓI and ΓIII come from the r +1 points where the contours get pinched by colliding poles. (Such pinching would usually cause integrals to diverge, but here the 54 divergence is cancelled by one of the s-independent theta-functions in Υ31 .) Therefore, letting B∗31 (x, t; q) := ds Υ31 (s, x, t; q) , 2πis Γ∗ Z (5.22) (with the obvious relation BI + BII + BIII = 0), we find 31 P3r1 (t; q) = BI31 (x, t; q) x=qr = BIII (x, t; q) x=qr . (5.23) We can repeat the analysis for the unknot U = 01 and figure-eight knot 41 . The superpolynomials of these knots are given by [74, 72, 75, 38]: Sr r r 3 P 01 (a, q, t) = a− 2 q 2 (−t)− 2 r Sr P 41 (a, q, t) = (a(−t)3 ; q)r (q; q)r (5.24) r X 3 r r (a(−t)3 ; q)r (aq −1 (−t); q)k (aq r (−t)3 ; q)k a−k− 2 q 2 +k(1−r) (−t)−2k− 2 r , (5.25) (q; q)k (q; q)r−k k=0 and Poincaré polynomials for G = SU (2), i.e. specializations to a = q 2 , normalized by (−1)r , are given by 3 1 P0r1 (t; q) = (−q 2 )−r (−t)− 2 r P4r1 (t; q) = (q 2 (−t)3 ; q)r (q; q)r r X 3 1 (q 2 (−t)3 ; q)r (q(−t); q)k (q 2 q r (−t)3 ; q)k (−q 2 )−r−2k(1+r) (−t)−2k− 2 r , (q; q)k (q; q)r−k (5.26) (5.27) k=0 respectively. Repeating the above procedure, we find 1 P0r1 (t; q) = B 01 (x, t; q) x=qr , B 01 (x, t; q) := 3 −2 3 − θ− (1, −q 2 x(−t) 2 ) (q −1 /x)− /t )∞ ∞ (−q (5.28) 1 3 − − −1 −2 − θ (x, −q 2 (−t) 2 ) (q )∞ (−q /(xt3 ))∞ for the unknot, and 1 Υ41 (s, x, t; q) := 1 1 − θ− (−q 2 x, q 2 tx, (−t)− 2 ) − (−1/(q 2 t3 ))− ∞ (−1/(qt))∞ 2 θ (qs, t s) 1 1 − − − . 2 3 (s)− θ− (q, t2 , q 2 t, x(−t)− 2 ) ∞ (−1/(qts))∞ (x/s)∞ (−1/(q xt s))∞ (5.29) for the figure-eight knot. In the latter case, the integrand Υ41 has four half-lines of poles in the s− − 2 3 − plane, coming from the four factors (s)− ∞ , (−1/(qts))∞ , (x/s)∞ , (−1/(q xt s))∞ in the denominator of (5.29). Let ΓI , ΓII , ΓIII , ΓIV be contours encircling these respective half-lines of poles. A formal 55 sum of residues along poles in the first half-line, evaluated at x = q r , most directly gives P4r1 (t; q); but the actual integral along ΓI does not converge for generic x. In contrast, the integrals along ΓII , ΓIII , ΓIV always converge, and 41 P4r1 (t; q) = BIII (x, t; q) x=qr = “BI41 (x, t; q) x=qr ” , (5.30) where B∗41 (x, t; q) := ds Υ41 (s, x, t; q) . Γ∗ 2πis Z (5.31) These examples indicate how the analysis may be extended to other knots and links (for example, those whose superpolynomials are found in [56, 57]), and to Poincaré polynomials of other homological invariants. In general, the required integrals will not be one-dimensional, but will require higher-dimensional integration cycles. Generalizations of some results in this chapter to other knots and links are also discussed in section 5.5.4. 5.4 Knot polynomials as partition functions of T [M3 ] In this section, we construct some examples of 3d N = 2 theories T [M3 ] for knot complements M3 (and G = SU (2)) with the properties outlined above. In particular, we would like the vacua of T [M3 ] on R2 × S 1 to match all flat connections on M3 . Although our strategy will be a little indirect, it is based on a simple key observation: the contour integral (5.10) for colored Poincaré polynomials has the form of localization integrals in supersymmetric 3d N = 2 theories as well as in Chern-Simons theory on certain 3-manifolds. Indeed, powerful localization techniques reduce the computation of Chern-Simons partition functions to finite dimensional integrals of the form (5.10), where the choice of the contour is related to the choice of the classical vacuum [76, 77, 78, 79, 80], as we briefly reviewed in section 5.6. Similar — and, in fact, closer to our immediate interest — contour integrals of the form (5.10) appear as a result of localization in supersymmetric partition functions of 3d N = 2 theories, such as the (squashed) sphere partition function [15, 18], the index [19, 20, 21], and the vortex partition function [39] or the half-index [25]. Since in the last case the space-time is non-compact it requires a choice of the asymptotic boundary condition or vacuum of the theory on R2 ×q S 1 , which manifests as a choice of the integration contour in the localization calculation. (The integrand is completely determined by the Lagrangian of 3d N = 2 theory.) This has to be compared with the first two cases, where localization of 3-sphere partition function and index lead to a contour integral with canonical choices of the integration contour. Therefore, in order to interpret (5.10) as a suitable partition function of 3d N = 2 theory in this thesis we mainly focus on half-indices, vortex partition functions, and their IR variants 56 called holomorphic blocks [13] that do not necessarily come from localization. This gives us enough flexibility to interpret (5.10) and we generically expect that the full set of blocks for T [M3 ], labelled by a full set of vacua, corresponds to a complete basis of independent convergent contours for the integrals of section 5.3. On the other hand, we also expect that a basis of convergent contours is in 1–1 correspondence with flat connections on M3 : vacua of T [M3 ] ↔ hol’c blocks ↔ convergent contours ↔ flat conn’s . (5.32) The reason to expect these correspondences to hold was discussed in the case without U (1)t in [39, 25, 13]. Similar idea is expected to hold with U (1)t symmetry and is outlined more carefully in section 5.4.1. In order to capture all flat connections, it turns out to be crucial that we start with Poincaré polynomials for knot homology rather than unrefined Jones polynomials. In section 5.4.2 we then demonstrate the construction of T [M3 ] in a few examples. In section 5.4.3 we examine the physical meaning of the limit x → q r that recovers Poincaré polynomials from T [M3 ]. We argue that it is a combination of Higgsing and creation of a line operator in T [M3 ], as on the left-hand side of (5.8). We also show that Poincaré polynomials can be obtained by directly taking residues of S 2 ×q S 1 indices and Sb3 partition functions of T [M3 ], bypassing holomorphic blocks. 5.4.1 Recursion relations for Poincaré polynomials One understanding of why contour integrals as in section 5.3 should capture all flat connections on a knot complement follows from looking at the q-difference relations that the integrals satisfy. r (t; q) for colored SU (2) knot homology of a knot Let us start with the Poincaré polynomials PK K. As found in [37, 38, 56], the sequence of Poincaré polynomials obeys a recursion relation of the form r bref (b A x, yb; t; q) · PK (t; q) = 0 , (5.33) r+1 4 r r r bref (b where A x, yb; t; q) is a polynomial operator in which x b, yb act as x bPK = q r PK and ybPK = PK . bref (b The limit q → 1 of A x, yb; t; q) is a classical polynomial Aref (x, y; t), whose subsequent t → −1 limit contains the classical A-polynomial of K [34] as a factor, q→1 t→−1 bref (b A x, yb; t; q) → Aref (x, y; t) → A(x, y) . (5.35) The physical interpretation of the classical A-polynomial A(x, y) goes back to [27]. Its roots at fixed 4 As mentioned before, notation l and m for longitude and meridian eigenvalues are related to y and x as l ↔ y, and similarly at the level of operator. m2 ↔ x (5.34) 57 x are in 1-1 correspondence with all flat connections on M (with fixed boundary conditions at K), but the root corresponding to the abelian flat connection is distinguished because it comes from a universal factor (y − 1) in A(x, y) as discussed in section 3.2. However, the t-deformed polynomial Aref (x, y; t) is irreducible (at least in simple examples5 ) in the sense that (y − 1) is not factorized, and none of its roots is more or less important than the others. Alternatively, note that the t → −1 limit of Aref (b x, yb; t; q) leads to a shift operator known as the b x, yb; q), which annihilates colored Jones polynomials [27, 81] as discussed quantum A-polynomial, A(b in section 3.2. One can also consider a–deformations of these shift operators. Such a deformation of the quantum A-polynomial was called Q-deformed A-polynomial in [82], and it agrees with the mathematically defined augmentation polynomial of [83, 84]. More generally, one can consider shift bsuper (b operators A x, yb; a; t; q) depending on both a and t, which annihilate colored superpolynomials, and which were called super-A-polynomials in [38] (for a concise review see [85]). However, as mentioned above, we are only interested here in a = q 2 specializations. Now, in section 5.3 we expressed r PK (t; q) = Z ds = BP (x, t; q) x=qr ΥK (s, x, t; q) ΓP 2πis x=q r (5.36) for a suitable integrand ΥK and a choice of integration contour ΓP . It is easy to see that BP (x, t; q) satisfies a q-difference equation bref (b A x, yb; t; q) · BP (x, t; q) = 0 (5.37) even before setting x = q r , with x b, yb acting as x bBP (x, ...) = xBP (x, ...) and ybBP (x, ...) = BP (qx, ...). R More so, the integral Bα = Γα ds/s ΥK for any convergent integration contour Γα (that stays bref · B = 0, sufficiently far away from poles) should provide a solution to the q-difference equation A and one generally expects that a maximal independent set of integration contours generates the full vector space of solutions.6 If we fix the values of x, t, and q, the convergent integration cycles Γα can be labelled by the roots y (α) (x, t) of the classical equation Aref (x, y; t) = 0 — i.e. by the flat connections on M3 with boundary conditions (meridian holonomy) fixed by x. The correspondence follows roughly by identifying the solutions to Aref (x, y; t) = 0 with critical points of the integrand ΥK (s, x, t; q) at q ≈ 1, then using downward gradient flow with respect to log |ΥK (s, x, t; q)| to extend the critical 5 To be more precise, both Aref (x, y; t) and A(x, y), obtained as appropriate limits of super-A-polynomials, may contain some additional factors. As explained in [29, 36] (for t = −1) and [37, 38] (for general t), these factors are necessary for quantization but are not associated to classical flat connections. For knots considered in this thesis these factors are independent of y, they do not affect the structure of roots of Aref (x, y; t) or A(x, y) at generic fixed x, and therefore they do not modify our discussion here. 6 This is technically a vector space over modular elliptic functions of (x, q), on which q-shifts acts trivially, as discussed in greater detail in [13]. 58 points into integration cycles Γα . bref · B = 0 as a contour integral (5.36), We have claimed that by writing one solution of A we can actually reproduce all other solutions from integrals on a full basis of contours Γα . This bref (and hence the classical Aref ) reasoning relies on an important assumption: that the quantum A is irreducible. Otherwise, we may only get solutions corresponding to one irreducible component. For this reason, it is crucial that we use refined knot polynomials and recursion relations rather than Jones polynomials and the quantum A-polynomial. See [86, 85] for further details as well as pedagogical introduction. To complete the chain of correspondences (5.32), we simply use [52, 51, 87, 11, 39, 46, 37, 56, 13, 88, 45] to translate the above observations to the language of gauge theory. Momentarily we R will engineer gauge theories T [M3 ] for which the integrals ∗ ds/s ΥK compute various partition bref and labelled by vacua of T [M3 ] on R2 × S 1 , i.e. classical functions on R2 × S 1 annihilated by A solutions of A(x, y; t) = 0. 5.4.2 3d N = 2 gauge theories for unknot, trefoil knot, and figure-eight knot Having rewritten the Poincaré polynomials of colored SU (2) knot homologies as special values of a contour integral, we try to engineer T [M3 ] so that the contour integral computes its partition function. In particular, by examining the integrand ΥK and from discussion in Chapter 2 and 4 we associate fugacities x, (−t), q flavor and R symmetries fugacity s U (1)s gauge symmetry (∗)− ∞ factors chiral multiplets θ− functions (mixed) Chern-Simons couplings (5.38) Then we can construct a putative UV description for T [M3 ] as an abelian Chern-Simons-matter theory. This approach is almost successful, and good enough for our present purposes, though we should mention an important caveat. In general, one must also specify relevant superpotential couplings for a UV description of T [M3 ], which are crucial for attaining the right superconformal theory in the IR; but it is very difficult to specify such couplings just by looking at partition functions. At the very least one would like to find superpotential couplings that break all “extraneous” flavor symmetries whose fugacities don’t appear in supersymmetric partition functions, and are not expected for the true T [M3 ]. Even this is difficult, because the naive prescription (5.38) leads to theories that simply don’t have chiral operators charged only under the extraneous symmetries. This problem was briefly discussed in Chapter 4, and solved by finding “resolved” theories with the same partition functions 59 as the naive ones, but with all necessary symmetry-breaking operators present [39, Section 4]. We also note that, while it is possible to construct the space of holomorphic blocks for a number of examples considered here, the construction is not always systematic, even in the simplest 3d N = 2 theories such as pure super-Chern-Simons theory. For this reason, it may be more convenient to work with other partition functions of theories T [M3 ] that include half-indices, i.e. UV counterparts of holomorphic blocks that are labeled by boundary conditions on R2 ×S 1 . A large class of boundary conditions in 3d N = 2 theories comes from 4-manifolds and will be discussed in section 5.6. It is expected that all holomorphic blocks can be reproduced (via RG flow) from suitable choice of boundary conditions in the UV. Other prominent examples of partition functions include the index (or, S 2 ×q S 1 partition function) [19, 20, 21] and the Sb3 partition function [15, 18], both of which will be discussed in section 5.4.3. Presently, we will follow the naive approach to obtain simple UV descriptions for putative T [M3 ]’s, where some but not all symmetry-breaking superpotential couplings are present. We expect that these theories are limits of the “true” superconformal knot-complement theories T [M3 ], where some marginal couplings have been sent to infinity. Thus, any observables of T [M3 ] that are insensitive to marginal deformations — such as holomorphic blocks, supersymmetric indices, massive vacua on S 1 , etc. — can be calculated just as well in our naive descriptions as in the true theories, as long as masses or fugacities corresponding to extraneous flavor symmetries are turned off by hand. This is sufficient for testing many of the properties we are interested in. Theory for unknot, T [01 ] The theory for the unknot that gives (5.28) was already discussed in [38] and has four chirals −2 −2 3 − /(xt3 ))− /t )∞ , (q −1 )− Φi , corresponding to the terms (q −1 /x)− ∞ . Letting x and (−t) ∞ , (q ∞ , (q be fugacities for flavor symmetries U (1)x and U (1)t , we use the rules discussed in section 4 of [52, 51, 87, 11, 39, 46, 37, 56, 13, 88, 45] to read off the precise charge assignments and levels of (mixed) background Chern-Simons couplings T [01 ] : Φ1 Φ2 Φ3 Φ4 U (1)x U (1)t U (1)R U (1)x −1 0 0 1 U (1)x 0 0 0 U (1)t 0 −3 0 3 U (1)t 0 0 0 U (1)R 0 −2 2 4 U (1)R 0 0 0 CS: (5.39) 60 (Here all background Chern-Simons couplings simply vanish.7 ). In this case, we can add an obvious superpotential W01 = µ Φ1 Φ2 Φ4 (5.40) that breaks most extraneous flavor symmetries and preserves U (1)x , U (1)t , and U (1)R (note that the operator in (5.40) has R-charge two). The chiral Φ3 is completely decoupled from the rest of the theory and rotated by an extraneous U (1) symmetry. We could break this U (1) by adding Φ3 to the superpotential (5.40), but prefer not to do this as it would forbid T [01 ] from having a supersymmetric vacuum. Ignoring the Φ3 sector, the putative unknot theory looks just like the 3d N = 2 XYZ model. Similarly, if we follow [37] and compactify T [01 ] on a circle turning on masses (i.e. complexified scalars in background gauge multiplets) for U (1)x and U (1)t , we find that the theory is governed by an effective twisted superpotential f0 = Li2 (x) + Li2 (−t3 ) + Li2 (−x−1 t−3 ) + 1 (log x)2 + 3 log x log(−t) + 9(log(−t))2 . W 1 2 (5.41) f0 an infinite contribution from the massless Φ3 ; this could be regularized (We have removed from W 1 by turning on a mass for the U (1) symmetry rotating Φ3 .) The equation for the supersymmetric parameter space,8 f ∂ W01 exp x = −y , ∂x (5.42) becomes the refined A-polynomial equation 3 (−t) 2 y = 1 + t3 x , 1−x (5.43) which further reduces to the unknot A-polynomial y − 1 = 0 at t → −1. Equation (5.43) has a unique solution in y at generic fixed x, t, corresponding to the unique, abelian flat connection on the unknot complement (with fixed holonomy eigenvalue x on a cycle linking the unknot). 7 We can multiply an extra normalization factor to SU (2) Poincaré polynomials to make the mixed IR CS levels for U (1)t to be integers, but we will work formally without such an extra normalization. 8 On the RHS we define an effective FI parameter as −y rather than y in order to match the knot-theoretic A-polynomial below. This is correlated with the renormalization of Poincaré polynomials above by (−1)r . 61 Theory for trefoil knot, T [31 ] In this case, the integrand (5.20) suggests a theory with six chirals, with charges and Chern-Simons levels T [31 ] : U (1)s U (1)x U (1)t U (1)R U (1)s −1/2 3/2 5/2 5/2 U (1)x 3/2 0 0 2 −3 U (1)t 5/2 0 0 0 −2 U (1)R 5/2 2 0 (5.44) 0 Φ1 Φ2 Φ3 Φ4 Φ5 Φ6 V− U (1)s −1 1 1 0 0 0 0 U (1)x 0 −1 0 0 0 1 −1 , U (1)t 0 0 1 −1 −3 3 U (1)R 0 0 2 0 −2 4 CS : This is now a gauge symmetry with a dynamical U (1)s symmetry in addition to U (1)x and U (1)t flavor symmetries. Standard analysis of [2] shows that this theory has a gauge-invariant anti-monopole operator V− formed from the dual photon, with charges as indicated in the table. Altogether we can write a superpotential W31 = µ1 Φ1 Φ2 Φ5 Φ6 + µ2 Φ1 Φ3 Φ4 + µ3 Φ6 V− (5.45) that preserves all symmetries we want to keep, and breaks almost all other flavor symmetry. There remains a single extraneous U (1), just like in the unknot theory, which plays (roughly) the role of a topological symmetry dual to U (1)s . When compactifying the theory on a circle with generic twisted masses x and (−t) for U (1)x and U (1)t , and scalar s in the U (1)s gauge multiplet, we obtain the effective twisted superpotential f3 = Li2 (s) + Li2 (−1/(st)) + Li2 (x/s) + Li2 (−t3 ) + Li2 (−t) + Li2 (−1/(t3 x)) W 1 + 21 (log s)2 + log s(6 log t + 2 log x) + log x(log x + 3 log(−t)) + 10(log t) (5.46) 2 . f3 /∂s) = 1, namely The critical-point equation exp s ∂ W 1 t2 (1 + st)(s − x)x =1 s(1 − s) (5.47) determines two solutions in s at generic values of x and t; plugging these into the SUSY-parameterspace equation f3 /∂x) = −s2 (−t)−3/2 −y = exp x ∂ W 1 1 + t3 x s−x (5.48) then determines two values of y. More directly, they are solutions of the quadratic 1 2 2 2 2 2 3 2 5 3 6 4 3 3 3 4 2 Aref 31 (x, y; t) = (1 − x)t y − (1 − t x + 2t x + 2t x + t x + t x )(−t) y + t (x + t x ) = 0 , . 62 3 which collapses to the Aref 31 (x, y; −1) = (x − 1)(y − 1)(y + x ), the trefoil’s A-polynomial (with an extra (x − 1) factor) as t → −1. Thus T [31 ] has vacua corresponding to both of the flat SL(2, C) connections on the trefoil complement, one irreducible, and one abelian. The two independent holomorphic blocks BII and BIII of (5.22) (or, more precisely, some linear combinations of these blocks) are in 1-1 correspondence with the two flat connections. Theory for figure-eight knot, T [41 ] Finally, for the figure-eight knot, the integrand (5.29) suggests a theory with U (1)s gauge symmetry, U (1)x × U (1)t flavor symmetry, and six chirals of charges T [41 ] : Φ1 Φ2 Φ3 Φ4 Φ5 Φ6 U (1)s −1 1 1 0 1 0 U (1)x 0 −1 0 0 1 0 U (1)t 0 0 1 −1 3 −3 U (1)R 0 0 2 0 4 −2 (5.49) The net Chern-Simons couplings all turn out to vanish. This particular theory does not admit gauge-invariant monopole or anti-monopole operators. We can introduce a superpotential W41 = µ1 Φ1 Φ3 Φ4 + µ2 Φ21 Φ2 Φ5 Φ6 , (5.50) which breaks flavor symmetry to U (1)4 , including U (1)x × U (1)t . Thus there are two extraneous U (1)’s, including the topological symmetry of the theory. As before, we can find an effective twisted superpotential on R2 × S 1 of the form f4 = Li2 (s) + Li2 (x/s) + Li2 (−1/(st)) + Li2 (−1/(sxt3 )) + Li2 (−t) + Li2 (−t3 ) + log’s , W 1 (5.51) whose critical point equation f4 ∂W 1 exp s ∂s = (1 + st)(s − x)(1 + st3 x) =1 (1 − s)st2 x (5.52) generically has three solutions in s — which in turn determine 3 f4 /∂x ∼ 1 + st x . y = − exp x ∂ W 1 s−x (5.53) 63 More directly, the solutions in y are roots of the cubic 9 3 2 2 2 2 3 2 4 3 5 3 5 4 6 4 7 5 2 2 3 Aref 41 = (x − x )(−t) y − (1 + tx − t x + 2t x + 2t x + 2t x + 2t x − t x + t x + t x )ty 1 + (−1 − tx + t2 x − 2t3 x2 − 2t4 x2 + 2t4 x3 + 2t5 x3 − t6 x4 + t7 x4 + t8 x5 )(−t) 2 y − (x2 + t3 x)t3 , 2 which deforms the standard figure-eight A-polynomial Aref 41 (x, y; t = −1) = (x−1)(y−1)(x −(1−x− 2x2 −x3 +x4 )y +x2 y 2 ) . Thus T [41 ] has massive vacua on S 1 corresponding to all three flat SL(2, C) connections on the figure-eight knot complement, two irreducible and one abelian. Again, these flat 41 41 41 connections label linear combinations of the three independent holomorphic blocks BII , BIII , BIV in (5.31). 5.4.3 Vortices in S 2 ×q S 1 and Sb3 Having obtained a theory T [M3 ] whose vacua on R2 × S 1 match flat connections on the knot complement M3 , it is interesting to probe its other protected observables. Here we focus on the S 2 ×q S 1 indices of T [M3 ], and make some preliminary observations as to the nature of the “Poincaré polynomial theories” Tpoly [M3 ; r] on the left-hand side of the flow diagram (5.8). The 3d index [19, 20, 21] of a knot-complement theory, or equivalently a partition function on 2 S ×q S 1 , depends on three fugacities q, ξ, τ and two integer monopole numbers n, p : fugacity monopole # symmetry q − combo of U (1)J ⊂ SO(3)Lorentz and U (1)R ξ n U (1)x τ p U (1)t (5.54) R We’ll consider “twisted” indices I(ζ, n; τ, p; q) = Tr Hn,p (S 2 ) eiπR q 2 −J ζ ex τ ep as in [25, 13], in which case it’s convenient to regroup fugacities into pairs of holomorphic and anti-holomorphic variables q = q, qe = q −1 ; n x = q2ξ, n x e = q 2 ξ −1 ; p −t = q 2 τ , p −e t = q 2 τ −1 . (5.55) r Then we find in examples below that the indices I[M3 ] of T [M3 ] develop poles at n = r and ξ = q 2 , or (x, x e) = (q r , 1), whose (logarithmic) residue is the r-th Poincaré polynomial of the colored SU (2) knot homology, Res(x,∼ I[M3 ] = x )→(q r ,1) r lim (1 − q 2 ξ −1 ) · I[M3 ](ξ, n; τ, p; q) ξ→q r/2 n=r r = PK (t; q) . (5.56) A similar statement holds for Sb3 partition functions. The Sb3 partition function Zb [15, 18] of a knot-complement theory depends on the ellipsoid deformation b as well as two dimensionless 64 complexified masses mx , mt for U (1)x , U (1)t , which are conveniently grouped into holomorphic and anti-holomorphic parameters 2 2 q = e2πib , qe = e2πi/b ; x = e2πbmx , x e = e2πmx /b ; −t = e2πbmt , −e t = e2πmt /b . (5.57) Then the Sb3 partition function has poles at mx = ibr, or (x, x e) = (q r , 1), with Res(x,∼ Z [M3 ] = x )→(q r ,1) b r lim (mx − ibr) · Zb [M3 ](mx , mt ; b) = PK (t; q) . mx →ibr (5.58) As discussed in section 2.4, both I[M3 ] and Zb [M3 ] take the form of a sum of products of holomorphic blocks [13] , I[M3 ], Zb [M3 ] ∼ X eα (e Bα (x, t; q)B x, e t; qe) , (5.59) α and our theory T [M3 ] was engineered so that the x → q r specialization of a specific linear combination of blocks BP would reproduce Poincaré polynomials. Below we will choose a convenient basis of blocks so that BP is one of the Bα ’s, and manifestly gives the only contribution to the residues (5.56), (5.58). (Nevertheless, in the natural basis of blocks labelled by flat connections at fixed (x, t ≈ −1, q = 1), BP may easily correspond to a sum over multiple flat connections, including the abelian one.) Taking the residue of a pole in an index such as (5.56) has an important physical interpretation, which was discussed in [59] in the context of 4d indices and, closer to our present subject, in [60, 61] in the context of 3d indices. Let us suppose that I[M3 ] is a superconformal index — i.e. that we have adjusted R-charges to take their superconformal values. Then the index counts chiral operators at the origin in R3 , and a pole signals the presence of an unconstrained operator O whose vev can parametrize a flat direction in the moduli space of T [M3 ]. Taking the residue of the pole is equivalent to giving a large vev to O, thus Higgsing any flavor symmetries under which O transforms, and decoupling massless excitations of T [M3 ] around this vev. Consider, for example, the pole at (x, x e) = (1, 1), or (ξ, n) = (1, 0). The pole suggests the presence of an operator Ox , of charge +1 under U (1)x , in the zero-th U (1)x monopole sector. The contribution of this operator and its powers to the index is (1 + ξ + ξ 2 + . . .) × I 0 = 1 × I0 . 1−ξ (5.60) Taking the residue I 0 amounts to giving a vev to Ox and decoupling massless excitations around it, thereby Higgsing U (1)x symmetry. One can interpret I 0 as the index of a new superconformal theory, the IR fixed point of a flow triggered by the vev hOx i. 65 r More generally, taking a residue at (x, x e) = (q r , 1) or (ξ, n) = (q 2 , r) gives a space-dependent vev (with nontrivial spin) to an operator in the r-th monopole sector. This not only Higgses the U (1)x symmetry of T [M3 ] but creates a vortex defect. We therefore expect that the residue of I[M3 ] at (x, x e) = (q r , 1) is the index of a new 3d theory Tpoly [M3 , r] in the presence of a (complicated) line operator. In the context of 4d theories T [C; G] coming from compactification of the 6d (2, 0) theory on a punctured Riemann surface C, taking the residue at a pole in the index amounted to removing a puncture from C — or more generally replacing the codimension-two defect at the puncture by a dimension-two defect in a finite-dimensional representation of G. Similarly, we expect here that taking a residue replaces the codimension-two defect along a knot K ⊂ M by a dimension-two defect in the (r +1)-dimensional representation of SU (2). We hope to elucidate this interpretation in future work. We proceed to examples of (5.56). Our conventions for indices follow [25, 13]. Below, all indices depend on fugacities from (5.55) as well as the pair k s = q2σ, k se = q 2 σ −1 , (5.61) which is used for summations/integrations. We assume |q| < 1, as is physically sensible for the index. Thus, the convergent q-Pochhammer symbols are (z)∞ := (z; q)∞ = Q∞ i i=1 (1 − q z) , (5.62) and theta-functions are θ(z1 , ..., zn ) := θ(z1 ; q) · · · θ(zn ; q) , 1 1 θ(z; q) := (−q 2 z)∞ (−q 2 z −1 )∞ . (5.63) Since we do our calculations while maintaining a manifest factorization into holomorphic blocks, results for Sb3 follow immediately from their index analogues, by reinterpreting the meaning of x e, e t, etc. 66 Unknot The index of the unknot theory T [01 ] from (5.39) is given equivalently by 1 3 3 I[01 ] = (−q 2 )n ξ 2 p τ 2 n 1 3 1 2 3 2 (5.64) 2 (−1/(qxt3 ))∞ θ(1, −q x(−t) ) (x−1 )∞ (−1/(qt3 ))∞ id θ(x, −q 2 (−t) 2 ) = = (q/e x)∞ (−q 2 /e t3 )∞ (−1/(qxt3 ))∞ (x−1 )∞ (−q −1 /t3 )∞ (−q 2 /(e xe t3 ))∞ 1 3 1 3 θ(x, −q 2 (−t) 2 , −q − 2 x e(−e t) 2 ) (q/e x)∞ (−q 2 /e t3 )∞ (−1/(qxt3 ))∞ × . 1 3 1 3 xe t3 ))∞ θ(e x, −q − 2 (−e t) 2 , −q 2 x(−t) 2 ) (x−1 )∞ (−q −1 /t3 )∞ (−q 2 /(e In the first line, we simply write down the index as defined by the theory — with the massless chiral Φ3 decoupled in order to remove an otherwise infinite factor. In the second line, we show that 2 this index comes from a fusion norm B 01 (x, t; q) id of the holomorphic block (5.28), with (q −1 )− ∞ removed. Since we are working at |q| < 1, we replace all q-Pochhammer symbols and theta-functions (z)− ∞ → 1 , (q −1 z)∞ θ− (z) → 1 θ(z) (5.65) in the definition of the block. In the third line, we explicitly write out what the fusion product means, following [13]. r We could take the limit (ξ, n) → (q 2 , r) in the first line of (5.64); after setting n = r, we would r find a pole at ξ → q 2 whose residue is the Poincare polynomial PUr (t, q). But it is more illustrative to take the equivalent limit (x, x e) → (q r , 1) in the factorized expression on the last line. Setting x e = 1 produces no divergence. The pole we are looking for comes from (x−1 )∞ in the denominator. We get lim ∼ (1 − q −r x) I[U ] = (x, x)→(q r ,1) 1 3 1 3 θ(q r , −q 2 (−t) 2 , −q − 2 (−e t) 2 ) (q)∞ (−q 2 /e t3 )∞ (−q −r−1 /t3 )∞ × 1 3 1 3 t3 )∞ θ(1, −q − 2 (−e t) 2 , −q 2 +r (−t) 2 ) (q −1 ; q −1 )r (q)∞ (−q −1 /t3 )∞ (−q 2 /e 1 3r = (−q 2 )−r (−t)− 2 (−q 2 t3 )r = PUr (t; q) . (q)r (5.66) Note how the e t dependence completely cancelled out of the problem. If we had taken a more general 0 0 limit (x, x e) → (q r , q r ), we would have found a similar pole, with residue PUr (t; q)PUr (e t; q −1 ). The 0 fact that the e t dependence cancels out follows from the simple identity PUr =0 (e t; q −1 ) = 1. Trefoil For the trefoil, the theory T [31 ] of (5.44) leads to an integral formula for the index, I[31 ] = I0 XI k∈Z 3 dσ θ(−q − 2 sex e(−e t)3 ) (qs)∞ (1/(−st))∞ (qx/s)∞ , 3 2πiσ θ(−q 2 sx(−t)3 ) (e s)∞ (q/(−e se t))∞ (e x/e s)∞ (5.67) 67 where I0 = 1 3 3 3 3 3 θ(−q − 2 x (−1/(qxt3 ))∞ (−q 2 /e t3 )∞ (−q/e t)∞ e, −q − 2 x e(−e t) 2 , −q 2 x(−t)3 , −q 2 (−t) 2 , x) × 1 3 3 3 3 3 − − 2 3 3 3 e e e 2 2 2 2 2 2 (−q /(e x t )) (−1/(qt )) (−1/t) θ(−q x, −q x(−t) , −q x e(−t) , −q (−t) , x e) ∞ ∞ ∞ (5.68) Again, we have chosen to regroup Chern-Simons contributions into ratios of theta-functions, separating out the x and x e dependence. The integrand in (5.67) has three pairs of half-lines of zeroes and poles in the σ-plane, coming from the three terms ( )∞ /( )∞ . They lie at I (qs)∞ /(e s)∞ II (−1/st)∞ /(−q/e se t)∞ III (qx/s)∞ /(e x/e s)∞ − k+p 2 +m σ = q − 2 +1+m ξ k−n k σ=q σ = q 2 +m k σ = q 2 −1−m τ −1 σ = q 2 −m ξ m ≥ max(−k, 0) m ≥ max(k + p, 0) m ≥ max(k − n, 0) zeroes σ = q − 2 −1−m poles τ −1 k+p k−n (5.69) The real, physical contour in (5.67) should lie on or around the unit circle, separating each half-line of zeroes from its corresponding half-line of poles. We also observe that the integrand of (5.67) vanishes as |σ| → ∞, if we stay away from half-lines of poles. Thus we can attempt to deform the contour outwards, closing it around σ = ∞. We pick up the poles in lines II and III, obtaining an expression of the form I[31 ] = I0 ||BII ||2id + ||BIII ||2id , (5.70) where9 ||BII ||2id = ||BIII ||2id = X θ(−q − 21 +m x ee t2 ) k,m≥0 θ(−q 1 2 −k (−q −k t−1 )∞ (−q 2+k tx)∞ , t−1 )∞ (−q −1−m e tx e)∞ xt2 ) (q)k (q −1 ; q −1 )m (−q m+1 e X θ(q − 32 +m x e2 e t3 ) k,m≥0 θ(q 3 2 −k x2 t3 ) (q)k 1 1 (q −1 ; q −1 ) (q 1−k x)∞ (−q k /(xt))∞ . e)∞ (−q 1−m /(e xe t))∞ m (q m x (5.71a) (5.71b) The holomorphic blocks BII and BIII here correspond to integrals along contours ΓII and ΓIII in Figure 5.2, with substitutions of the form (x)− ∞ → 1/(qx)∞ to account for |q| < 1. Now, if we send (x, x e) → (q r , 1), the leading pole in line I can collide with the leading pole in line III, pinching the integration contour in the σ-plane, and leading to a divergence of the the index. We see this explicitly in the evaluated expression (5.70): while the prefactor I0 and the blocks ||BII ||2id are finite in this limit, the blocks ||BIII ||2id have the expected divergence. It comes from the denominator (q m x e)∞ in (5.71b), and occurs only for m = 0. The related factor (q 1−k x)∞ 9 A redefinition of summation indices turns the sum over k ∈ Z into sums k ≥ 0. 68 in the numerator vanishes as x = q r unless k ≤ r. Therefore, we find a residue ∼ (1 − x e) I[31 ] = lim (x, x)→(q r ,1) ∼ lim (x, x)→(q r ,1) (1 − x e) I0 ||BIII ||2id r = I0 (x = q , x e = 1; t, e t; q) 3 r X θ(q − 2 e t3 ) (q 1−k+r )∞ (−q k−r t−1 )∞ 3 k=0 θ(q 2 −k t3 ) (q)k (q)∞ (−q/e t)∞ = P3r1 (t; q) P301 (e t; q −1 ) = P3r1 (t; q) , (5.72) reproducing the superpolynomial after some straightforward manipulations. Figure-eight knot The setup for the figure-eight knot is almost identical to that for the trefoil. Now the index is given by I[41 ] = I0 with I0 = XI k∈Z t2 se) (qs)∞ (−1/(ts))∞ (qx/s)∞ (−1/(qxt3 s))∞ dσ θ(q −1 se, e , 2πiσ θ(qs, t2 s) (e s)∞ (−q/(e tse))∞ (e x/e s)∞ (−q 2 /(e xe t3 se))∞ 1 1 1 1 1 θ(t2 , q 2 t, x(−t)− 2 , −q − 2 x e, q − 2 e t3 )∞ (−q/e t)∞ tx e, (−e t)− 2 ) (−q 2 /e × . 1 1 1 1 1 3 − − − 2 (−1/qt )∞ (−1/t)∞ θ(e t ,q 2e t, x e(−e t) 2 , −q 2 x, q 2 tx, (−t) 2 ) (5.73) (5.74) There are four pairs of half-lines of zeroes and poles in the integrand; three are identical to those in the trefoil integrand above, which we denote I, II, III as in (5.69), and there is one new pair IV : zeroes σ = q − poles σ = q k+n+3p −1+s 2 k+n+3p −2+s 2 ξ −1 τ −3 ξ −1 τ −3 , for m ≥ max(k + n + 3p, 0) . (5.75) We close the contour around σ = ∞ (where the integrand generically vanishes), picking up the poles in lines II, III, and IV, to give I[41 ] = I0 ||BII ||2id + ||BIII ||2id + ||BIV ||2id , (5.76) 69 with ||BII ||2id = X k,m≥0 θ(−q m /e t, −q m+1 e t) 1 (−q −k /t)∞ (−q 2+k xt)∞ (q k /(xt2 ))∞ , −k −k−1 −1 −1 θ(−q /t, −q t) (q)k (q , q )m (−q m+1 /e t)∞ (−q −m−1 x ee t)∞ (q 1−m /(e xe t2 ))∞ (5.77a) ||BIII ||2id = m−1 m e2 1−k k k−1 2 3 X θ(q x e, q t x 1 (q x)∞ (−q /(xt))∞ (−q /(x t ))∞ e) , 1−k −k 2 −1 −1 m 1−m 2−m e θ(q x, q t x) (q)k (q , q )m (q x t3 ))∞ e)∞ (−q /(tx e))∞ (−q /(e x2 e k,m≥0 (5.77b) −q m+1 −q m+2 ∼ X θ ∼∼3 , ∼ 1 (−q −k−1 /(xt3 ))∞ (q 2+k xt2 )∞ (−q k+3 x2 t3 )∞ xt 2 t x−k−1 , ||BIV ||id = −k−2 −1 −1 −q −q (q)k (q , q )m (−q m+2 /(e t3 )∞ xe t3 ))∞ (q −m−1 x ee t2 )∞ (−q −m−2 x e2 e , 3 k,m≥0 θ xt xt (5.77c) The holomorphic blocks in these expressions correspond to the integration cycles discussed above (5.31) (with the usual translation from |q| > 1 to |q| < 1). Now as (x, x e) → (q r , 1), the prefactor I0 along with ||BII ||2id and ||BV I ||2id all have finite limits; while ||BIII ||2id has a pole due 1/(q m x e)∞ at m = 0, and is nonvanishing for k ≤ r. As in the case of the trefoil, the divergence can be attributed to the poles of lines I and III pinching the contour of the integrand (5.73). We then find ∼ lim (x, x)→(q r ,1) (1 − x e) I[41 ] = ∼ lim (x, x)→(q r ,1) (1 − x e) I0 ||BIII ||2id = P4r1 (t; q) P401 (e t; q −1 ) = P4r1 (t; q) . 5.5 (5.78) The t = −1 limit and DGG theories Above, we saw that sending x → q r in partition functions of T [M3 ] (and perhaps discarding an overall divergence) produced finite Poincaré polynomials of colored SU (2) knot homologies. Once the Poincaré polynomials are obtained, we are free to send t → −1 to directly recover the colored Jones polynomials. No further divergences are encountered. Physically, we proposed an identification of the regularized x → q r limit with a physical “Higgsing” process, by which an operator in T [M3 ] charged under U (1)x is given a space-dependent vev, initiating an RG flow to a new theory in the presence of a line defect. Subsequently sending t → −1 should not correspond to any further flow. One may wonder what would happen if we sent t → −1 before x → q r . We present evidence in this section that this initiates a different RG flow in T [M3 ], which ends at a DGG theory TDGG [M3 ]. In particular, an operator Ot is given a (constant) vev, breaking the U (1)t symmetry characteristic of T [M3 ]. Moreover, vacua of T [M3 ] on R2 ×S 1 that correspond to abelian or reducible flat connections on M3 are lost. As above, our analysis will be largely example-driven. In section 5.5.1 we examine how the trefoil 70 and figure-eight knot theories of section 5.4.2 flow to DGG theories. We verify in section 5.5.2 that t → −1 limits induce divergences in S 2 ×q S 1 indices, indicative of Higgsing. Then in section 5.5.3 we use effective twisted superpotentials on R2 × S 1 to better understand how vacua corresponding to abelian flat connections decouple. 5.5.1 The DGG theories We can see an explicit example of the proposed DGG flow by considering the trefoil theory T [31 ] of (5.44). If we turn off the real mass for the flavor symmetry U (1)t , then the chiral operator Ot = Φ4 can get a vev, hΦ4 i = Λ . (5.79) The vev breaks U (1)t , but no other symmetries. Moreover, it induces a complex mass for Φ1 and Φ3 due to the superpotential W31 = µ1 Φ1 Φ2 Φ5 Φ6 + µ2 Λ Φ1 Φ3 + µ3 Φ6 V− . (5.80) Therefore, taking Λ → ∞, we may decouple fluctuations of Φ4 and integrate out Φ1 and Φ3 , arriving at T 0 [31 ] : Φ2 Φ5 Φ6 V− U (1)s 1 0 0 0 U (1)x −1 0 1 −1 U (1)R 0 −2 4 −2 , CS : U (1)s U (1)x U (1)R U (1)s −1/2 3/2 5/2 U (1)x 3/2 0 2 U (1)R 5/2 2 0 . (5.81) with superpotential W30 1 = µ03 Φ6 V− . (5.82) At this point, we observe that T [31 ] has a sector containing a U (1)s gauge theory with a single charged chiral Φ2 , together with minus half a unit of background Chern-Simons coupling. This sector can be dualized to an ungauged chiral ϕ as in [39, Sec 3.3], a consequence of a basic 3d mirror symmetry [89, 3]. Indeed, the dual ungauged chiral is identified with the (anti-)monopole operator ϕ = V− of U (1)s ! Thus, T 0 [31 ] is dual to 00 T [31 ] : Φ5 Φ6 ϕ U (1)s 0 0 0 U (1)x 0 1 −1 U (1)R −2 4 −2 , CS : U (1)x U (1)R U (1)x 3 6 U (1)R 6 ∗ . (5.83) 71 with W3001 = µ003 Φ6 ϕ. The superpotential lets us integrate out Φ6 and ϕ, leaving behind T 00 [31 ] TDGG [31 ] ⊗ TΦ5 . (5.84) Here Φ5 is a fully decoupled free chiral, while TDGG [31 ] is a slightly degenerate description of the DGG trefoil theory. Namely, TDGG [31 ] here is a “theory” consisting only of a background Chern-Simons coupling at level 3 for the flavor symmetry U (1)x , and some flavor-R contact terms given by the matrix on the RHS of (5.83). A similar “theory” was obtained by DGG methods in section 4.3 ([25, Section 4.3] for detail), using a degenerate triangulation of the trefoil knot complement into two ideal tetrahedra. It was interpreted as an extreme limit of the true DGG theory TDGG [31 ] in marginal parameter space. It is not surprising that we have hit such a limit, since, as discussed at the beginning of section 5.4.2, we are ignoring some marginal deformations. Our TDGG [31 ] becomes identical to that in section 4.3 upon shifting R-charges by minus two units of U (1)x charge. The shift is due to difference of conventions: we initially set x = q r in Poincaré polynomials whereas the equivalent choice for [39, 25] would be x = q r+1 . We can repeat this exercise for the figure-eight knot. The theory T [41 ] of (5.49) again has a chiral operator Ot = Φ4 that is charged only under U (1)t , and can get a vev when the real mass corresponding to U (1)t is turned off, hΦ4 i = Λ . (5.85) W41 = µ1 Λ Φ1 Φ3 + µ2 Φ21 Φ2 Φ5 Φ6 (5.86) Then the effective superpotential lets us integrate out Φ1 and Φ3 . We flow directly to a theory T [41 ] TDGG [41 ] ⊗ TΦ6 , (5.87) where Φ6 is a decoupled chiral and TDGG [41 ] : Φ2 Φ5 U (1)s 1 1 U (1)x −1 1 U (1)R 0 4 , (CS vanishing) (5.88) is basically the GLSM description of the CP1 sigma-model. It is equivalent (after shifting R-charges by minus two units of U (1)x charge) to the DGG theory in section 4.3 obtained from a triangulation 72 of the figure-eight knot complement into two tetrahedra.10 Again, this triangulation is a little degenerate as discussed in section 4.3 ([39, Section 4.6] for detail), so (5.88) should be viewed as a limit of the true TDGG [41 ], which has the same protected partition functions (index, half-indices, and holomorphic blocks). 5.5.2 Indices and residues The S 2 ×q S 1 indices of theories T [M3 ] help us to further illustrate the breaking of U (1)t by “Higgsing” and the flow to TDGG [M3 ]. As discussed in section 5.4.3, Higgsing corresponds to taking residues in an index. In particular, we expect here to find the indices IDGG [M3 ] of DGG theories as residues of I[M3 ] at (t, e t) → (−1, −1). Consider, for example, the index I[31 ] of the trefoil theory as given by (5.70). Sending t → −1, the prefactor I0 develops a pole due to the factor 1/(−1/t)∞ . This factor comes directly from the chiral Φ4 in T [31 ]. (The factor 1/(−1/(qt3 ))∞ in I0 , coming from the chiral Φ4 , also develops a pole, but it is not relevant for the Higgsing we want to do.) In addition, we see that ||BIII ||2id has a finite limit as (t, e t) → (−1, −1), whereas ||BIII ||2id vanishes due to (−q −k t−1 )∞ in the numerator. One way to understand this vanishing is to observe that the zeroes in line I of the index integrand perfectly cancel all poles in line II when (t, e t) = (−1, −1). Therefore, lim (1 − t)I[31 ] = ∼ t, t →−1 = = lim (1 − t) I0 ||BIII ||2id 1 “ (−q 2 /e t3 )∞ ” θ(−q − 2 x e, x)(1/(qx))∞ X (−1/(qt3 ))∞ (5.89) ∼ t, t →−1 1 2 θ(−q x, x e)(q 2 /e x)∞ 3 “ (−q 2 /e t3 )∞ ” θ(x, q − 2 x e2 ) 3 2 (−1/(qt3 ))∞ θ(e x, −q x2 ) “ (−q 2 /e t3 )∞ ” = IDGG [31 ] . (−1/(qt3 ))∞ = k,m≥0 1 3 1 e2 , −q 2 −k x) θ(−q m− 2 x (q)k (q −1 ; q −1 )m θ(−q 3 2 −k 1 x2 , −q m− 2 x e) “ (−q 2 /e t3 )∞ ” 3n 3n q ξ (−1/(qt3 ))∞ The resummation in the third line captures the duality between a charged chiral (Φ2 ) and a free chiral (ϕ = V− ) discussed in section 5.5.1. Then the expression q 3n ξ 3n matches the DGG trefoil index of [25], modulo a redefinition of R-charges ξ → q −1 ξ. The infinite prefactor (−q 2 /e t3 )∞ /(−1/(qt3 ))∞ → (q 2 )∞ /(q −1 )∞ is the contribution of the decoupled chiral Φ5 . When considering the t, e t → −1 limit of the figure-eight index I[41 ] from (5.76), the prefactor I0 has the same divergent term (−1/t)−1 ∞ that appeared for the trefoil. Moreover, the contribution ||BII ||2id to the figure-eight index vanishes, because poles of the index integrand in line II are cancelled 10 The equivalence is most directly seen using the polarization discussed in Appendix B and section 6.3 of [13]. The “degenerate” DGG theory for the figure-eight knot, a.k.a. the CP1 sigma-model, has three standard duality frames that are analyzed in section 5.1 of [13], and the most symmetric of these duality frames agrees with (5.88). Another frame matches section 4.6 of [39]. 73 by zeroes in line I. Thus, following a short calculation, the figure-eight index takes the form lim (1 − t)I[41 ] = ∼ t, t →−1 lim (1 − t) I0 ||BIII ||2id + ||BIV ||2id ∼ t, t →−1 (5.90) X (qx)k (q −1 x “ (−q 2 /e 1 t3 )∞ ” e)m (q k+1 (qx)2 )∞ 2n 2 )n (qξ) (−q = (−1/(qt3 ))∞ (q −1 ; q −1 )k (q)m (q −m (q −1 x e)2 )∞ k,m≥0 + (n, qξ) ↔ (−n, 1/(qξ)) = “ (−q 2 /e t3 )∞ ” IDGG [41 ] . (−1/(qt3 ))∞ We recognize in this the DGG index of the figure-eight knot, already split into two holomorphic blocks. For proper comparison to [25] or [13], we should again rescale ξ → q −1 ξ, or (x, x e) → (q −1 x, qe x). 5.5.3 Critical points and missing vacua We saw in section 5.5.2 that in the limit t, e t → −1, some parts of indices I[M3 ] vanished, while others contributed to IDGG [M3 ]. This is a reflection of the fact that the DGG theories TDGG [M3 ] don’t capture all information about flat connections on M3 , and in particular don’t have massive vacua on R2 × S 1 corresponding to abelian or reducible flat SL(2, C) connections. We can make this idea much more precise by considering the effective twisted superpotentials that govern theories T [M3 ] on R2 × S 1 . For example, for the trefoil, this was given by (5.46): f3 (s; x, t) = Li2 (s) + Li2 (−1/(st)) + Li2 (x/s) + Li2 (−t3 ) + Li2 (−t) + Li2 (−1/(t3 x)) W 1 (5.91) + 21 (log s)2 + log s(6 log t + 2 log x) + log x(log x + 3 log(−t)) + 10(log t)2 . It is important to note that this function on C∗ (parametrized by the dynamic variable s) has branch cuts coming from integrating out chiral matter that at some points in the s-plane becomes massless. In particular, each term Li2 (f (s)) has a cut along a half-line starting at the branch point f (s) = 1 and running to zero or infinity. Such cuts and their consequences have been discussed from various perspectives in e.g. [12, 10, 90, 62]. Often one writes the vacuum or critical-point equations as f3 /∂s = 1 , exp s ∂ W 1 (5.92) because in this form they are algebraic in s. However, when analyzing vacua of T [M3 ] on R2 × S 1 , f — and one must remember to lift solutions of (5.92) back to the cover of the s-plane defined by W to make sure they are actual critical points on some sheets of the cover. Now consider what happens if we send t → −1. The branch points of Li2 (s) and Li2 (−1/(st)), 74 located at s = 1 and s = −1/t, collide. (These branch points came directly from the chirals Φ1 and Φ3 , which we integrated out of T [31 ] in (5.81).) In the process, the half-line cuts originating at these branch points coalesce into a full cut running from s = 0 to s = ∞; this is easy to see from the inversion formula 2 Li2 (s) + Li2 (1/s) = − π6 − 12 log(−s)2 (s ∈ / [0, 1) ) . (5.93) Moreover, one of the solutions s∗ to (5.92), or rather its lift(s) to the covering of the s-plane, gets trapped between the colliding branch points and ceases to be a critical point as t → −1. One can see this from the explicit form of the critical-point equations (5.47), which are reduced from quadratic to linear order in s by a cancellation at t = −1. However, to properly interpret this limit, it is helpful to think about the branched cover of the s-plane as we have done. Physically, each solution of (5.92) is a vacuum of T [M3 ] on R2 × S 1 . As t → −1, the vacuum at s∗ is lost. This is possible precisely because the t → −1 limit is singular. Indeed, we know that t → −1 corresponds to making T [M3 ] massless, so that the reduction on R2 × S 1 is no longer fully f (s; x, t = −1) described by an effective twisted superpotential. The specialized superpotential W does not describe T [M3 ] itself at the massless point, but rather the Higgsed TDGG [M ] as found in section 5.5.1. In the case of the trefoil, the vacuum at s∗ close to t = −1 is labelled (via the 3d-3d correspon- dence) by the abelian flat connection on M3 = S 3 \K. Indeed, if we substitute the limiting t → −1 f /∂x = y, we find value of s∗ (namely s∗ = 1) into the SUSY-parameter-space equation exp x ∂ W f /∂x y∗ := exp x ∂ W =1 s ∗ at t = −1 , (5.94) corresponding to the abelian factor y − 1 = 0 of the trefoil’s classical A-polynomial. Thus we see explicitly that the DGG theory TDGG [31 ] loses a vacuum corresponding to the abelian flat connection. We may also perform this analysis at the level of holomorphic blocks. Holomorphic blocks are f — or more precisely by integration cycles Γα obtained labelled by (q-deformed) critical points of W f and approximately following gradient flow with respect to by starting at a critical point of W f . For the trefoil we can choose a basis of integration cycles given by ΓII and ΓIII in Figure Re log1 q W 5.2. The precise correspondence with critical points depends on x, t, q. Close to t = −1, however, it is clear that ΓII corresponds to the “abelian” critical point s∗ . As t → −1, the contour ΓII gets trapped crossing a full line of poles (resolutions of the classical branch cuts described above), and ceases to be a good holomorphic-block integration cycle in the sense of [13].11 Most importantly, it 11 Of course, Γ II is still a reasonable integration cycle, mathematically, at t = −1 and any finite q. The integral along it does reproduce a Jones polynomial as x → q r . It is tempting to wonder whether one could engineer such a 75 no longer flows from any classical critical point. Beautifully, the remaining contour ΓIII is isolated away from the point s∗ where half-lines of poles merge. The t → −1 limit of the corresponding block BIII (x, t; q) is precisely the holomorphic block of TDGG [31 ], labelled by the irreducible flat SL(2, C) connection, and contributing to the index (5.89). Analogous remarks apply to the figure-eight example. The 3d Higgsing and integrating out of Φ1 , Φ3 in T [41 ] translates on R2 × S 1 to branch points of Li2 (s) and Li2 (−1/(st)) colliding in (5.51), and trapping a critical point between them. Thus, as t → −1, T [41 ] looses one of its three massive vacua on R2 × S 1 — the one labeled by the abelian connection on the figure-eight knot complement. The TDGG [41 ] only has two massive vacua, labelled by irreducible flat connections. The remaining vacua correspond to the holomorphic blocks BIII and BIV , which at t → −1 become the holomorphic blocks of TDGG [41 ]. 5.5.4 Relation to colored differentials We expect that the Higgsing procedure found to relate T [M3 ] to TDGG [M3 ] in the examples above holds much more generally. We can actually recognize some key signatures of the reduction in a much larger family of examples, which include so-called thin knots. The phenomena described above follow from the structure of colored Poincaré polynomials for these knots. The structure of the Poincaré polynomials is highly constrained by the properties of colored differentials whose existence in S r -colored homologies was postulated in [69, 91], as well as by the so-called exponential growth. Using these properties, in [56] colored Poincaré polynomials of many thin knots, including the infinite series of (2, 2p + 1) torus knots and twist knots with 2n + 2 crossings, were uniquely determined. More precisely, colored differentials enable transitions between homology theories labeled by the r-th and k-th symmetric-power representations S r and S k . The existence of these differentials implies that Poincaré polynomials take the form of a summation (over k = 0, . . . , r), with the summand involving a factor (−aq −1 t; q)k . On the other hand, the exponential growth is the statement that for q = 1 (normalized) colored Poincaré polynomials (superpolynomials) satisfy the relation 1 r Sr S PK (a, q = 1, t) = PK (a, q = 1, t) . (5.95) If the uncolored superpolynomial on the right hand side is a sum of a few terms, its r’th power can be written as a (multiple) summation involving Newton binomials, which for arbitrary q turn out to be replaced by q-binomials [56, 57]. This structure can be clearly seen in the example of (2, 2p + 1) “Jones” cycle starting directly with TDGG [M3 ], with no prior knowledge of the full T [M3 ] — and what the physical meaning of this cycle might be. 76 torus knots considered in [56, 57], whose (normalized) colored superpolynomials take the form  r PTS2,2p+1 (a, q, t) = apr q −pr X  0≤kp ≤...≤k2 ≤k1 ≤r p r k1   k1 k2   kp−1 ··· × q (2r+1)(k1 +k2 +...+kp )−Σi=1 ki−1 ki t2(k1 +k2 +...+kp ) k1 Y kp  × (5.96) (1 + aq i−2 t). i=1 Here the last product originating from the structure of differentials, as well as a series of q-binomials originating from the exponential growth, are manifest (in this formula k0 = r). Poincaré polynomials for infinite families of twist knots derived in [56, 57] share analogous features. It becomes clear now that various properties of trefoil and figure-8 knots, discussed earlier, should also be present for other knots, such as thin knots discussed above. For example, as discussed in section 5.5.2, the divergence at t → −1 in the trefoil and figure-8 indices, I[31 ] and I[41 ], is a manifestation of a pole due to the factor 1/(−1/t)∞ . This factor originates from the q-Pochhammer symbol (−aq −1 t; q)k in corresponding Poincaré polynomials (5.13) and (5.25), after setting a = q 2 and rewriting this term in the denominator. As follows from the discussion above, such a factor is present in general for other thin knots (and represents the action of colored differentials), so for such knots an analogous pole at t → −1 should develop. We postulate that the residue at this pole in general reproduces indices IDGG [M ] for theories dual to other (thin) knots. Similarly, a decoupling of the abelian branch for more general knots is a consequence of the structure of superpolynomials described above. From this perspective, let us recall once more how this works for trefoil and figure-8 knot, just on the level of critical point equations (5.47) and (5.48), or (5.52) and (5.53). If we set t = −1 in (5.47) or (5.52), the ratio 1+st 1−s on the left hand side drops out of the equation (this is a manifestation of the cancelation (5.93) at the level of twisted superpotential ). In this ratio the numerator 1 + st has its origin in the (−aq −1 t; q)k term in superpolynomials (5.13) and (5.25), while the denominator 1 − s originates from q-Pochhamer (q; q)k being a part of the q-binomial in those superpolynomials. As explained above, such terms appear universally in superpolynomials for thin knots. Similarly, for t = −1 the equations (5.48) and (5.53) reduce to y = 1 (which represents the abelian branch that drops out when t → −1 is set first) due to a cancellation between the term in their numerator and s − x in denominator. The terms in numerator have the origin in (a(−t)3 ; q)r from unknot normalization (5.24), possibly combined with another term (aq r (−t)3 ; q)k representing colored differentials for figure-8 knot (5.25). The term s − x in denominator has its origin in (q; q)r−k ingredient of q-binomial. Analogous terms, responsible for cancellations, are also universally present in superpolynomials for other knots. The analysis is slightly more involved if Poincaré polynomials include multiple summations — e.g. for (2, 2p + 1) torus knots (5.96) — however one can check that similar cancellations between “universal” terms decrease the degree of saddle equations and result in the decoupling of the abelian branch. 77 5.6 Boundaries in three dimensions In this section we discuss the gluing along boundaries of M3 and the boundary conditions in 3d N = 2 theories T [M3 ]. In particular, understanding the operations of cutting and gluing M3 along a Riemann surface C opens a new window into the world of closed 3-manifolds. The basic idea of how such operations should manifest in 3d N = 2 theory T [M3 ] was already discussed e.g. in [51, 11] and will be reviewed below. The details, however, cannot work unless T [M3 ] accounts for all flat connections on M3 . This was recently emphasized in [58] where the general method of building T [M3 ] via gluing was carried out for certain homology spheres. After constructing 3d N = 2 theories T [M3 ] for certain homology spheres, we turn our attention to boundary conditions in such theories. Incorporating boundary conditions and domain walls in general 3d N = 2 theories was discussed in [62] and involves the contribution of the 2d index of the theory on the boundary / wall that is a “flavored” generalization of the elliptic genus. For theories of class R that come from 3-manifolds, many such boundary conditions come from 4-manifolds as illustrated in (5.9). In this case, the flavored elliptic genus of a boundary condition / domain wall is equal to the Vafa-Witten partition function of the corresponding 4-manifold [58]. 5.6.1 Cutting and gluing along boundaries of M3 It is believed that a 3-manifold with boundary C gives rise to a boundary condition in 4d N = 2 theory of class S, see Figure 2 in [39] or Figure 6 in [58]. This system can be understood as a result of 6d (2, 0) theory compactified on a 3-manifold with cylindrical end R+ × C and to some extent was studied previously.12 For example, when C = T 2 is a 2-torus (with puncture) the corresponding 4d N = 2 theory is actually N = 4 super-Yang Mills (resp. N = 2∗ theory). A simple class of 3-manifolds bounded by C includes handlebodies, which for a genus-g Riemann surface C is determined by a choice of g pairwise disjoint simple closed curves on C (that are contractible in the handlebody 3-manifold). For example, if C = T 2 , then the corresponding handlebody is a solid torus: M3 ∼ = S 1 × D2 . (5.97) It is labeled by a choice (p, q) of the 1-cycle that becomes contractible in M3 . In the basic case of (p, q) = (0, 1) the Chern-Simons path integral on M3 defines a state (in the Hilbert space HT 2 ) that is usually denoted |0i, so that we conclude |0i = |solid torusi (5.98) 12 See e.g. [51, 87, 11, 39, 92] for a sample of earlier work; unfortunately the methods of these papers cannot be used to recover all flat connections for general 3-manifolds, even in the simplest cases of knot complements. 78 It was proposed in [58] that the corresponding boundary condition in 4d theory T [C] is Nahm pole boundary condition [93, 94] that can be described by a system of D3-branes ending on D5-branes13 |0i = |Nahmi = |D5i (5.99) More generally, for M3 ∼ = S 1 × D2 obtained by filling in the cycle in homology class (p, q) the corresponding boundary condition is defined by a system of D3-branes ending on IIB five-branes of type (p, q). This class of boundary conditions can be easily generalized to other Riemann surfaces C and 3-manifolds with several boundary components. The latter correspond to domain walls in 4d N = 2 theories T [C], see e.g. [39, 58, 92] for details. For example, each element φ of the mapping class group of C corresponds, on the one hand, to a mapping cylinder M3 (with two boundary components identified via φ) and, on the other hand, to a duality wall of type φ in the 4d theory T [C]. In the case C = T 2 we have the familiar walls that correspond to the generators φ = S and φ = T of the SL(2, Z) duality group of N = 4 super-Yang-Mills, and the general “solid torus boundary condition” described above can be viewed as the IR limit of a concatenation of S- and T -walls with the basic Nahm pole boundary condition, see [58, pp.20-21] for details. For instance, S|0i = |Neumanni = |NS5i (5.100) Clearly, there are still many details to work out, but we have outlined the key elements necessary to glue 3-manifolds along a common boundary and, in particular, to illustrate why (5.5) must hold in a proper 3d N = 2 theory T [M3 ]. Suppose C = ±∂M3± is a common boundary component of 3-manifolds M3+ and M3− , which in general may have other boundary components, besides C. As we reviewed earlier, appropriately defined 3d N = 2 theories T [M3+ ] and T [M3− ] naturally couple to a 4d N = 2 theory T [C], which becomes dynamical upon the gluing process M3 = M3− ∪φ M3+ (5.101) Note, in the identification of the two boundaries here we included an element φ of the mapping class group of C that corresponds to duality wall in T [C]. Hence, the resulting theory T [M3 ] consists of a φ-duality wall in 4d N = 2 theory T [C] sandwiched between T [M3+ ] and T [M3− ]. At the level of partition functions, ZT [M3 ] = ZCS (M3 ) = hM3− |φ|M3+ i (5.102) A particularly simple and useful operation that involves (re)gluing solid tori a la (5.97)–(5.101) 13 Whether we identify the state |0i with D5 or NS5 is a matter of conventions. Here we follow the conventions of [58, 62]. 79 is called surgery. In fact, it is also the most general one in a sense that, according to a theorem of Lickorish and Wallace, every closed oriented 3-manifold can be represented by (integral) surgery along a link K ⊂ S 3 . Since the operation is defined in the same way on any component of the link L it suffices to explain it in the case when K has only one component, i.e. when K is a knot. Then, for a pair of relatively prime integers p, q ∈ Z, the result of q/p Dehn surgery along K is the 3-manifold: 3 Sq/p (K) := (S 3 − N (K)) ∪φ (S 1 × D2 ) (5.103) where N (K) is the tubular neighborhood of the knot, and S 1 × D2 is attached to its boundary by a diffeomorphism φ : S 1 × ∂D2 → ∂N (K) that takes the meridian µ of the knot to a curve in the homology class q[µ] + p[λ] (5.104) The ratio q/p ∈ Q ∪ {∞} is called the surgery coefficient. In what follows we discuss various aspects of cutting, gluing, and surgery operations. In particular, we shall see how the operations (5.102) and (5.103) manifest at various levels in 3d N = 2 theory T [M3 ] — at the level of SUSY vacua, at the level of twisted superpotential, and at the level of quantum partition functions — thereby illustrating the important role of abelian flat connections. Needless to say, there are many directions in which one could extend this analysis, e.g. to various classes of 3-manifolds not considered in this thesis, as well as more detailed analysis of the ones presented here, to higher rank groups G and to relation with known properties of homological knot invariants. 5.6.1.1 Compactification on S 1 and branes on the Hitchin moduli space A useful perspective on our 3d-4d system can be obtained by compactification on S 1 and studying the space of SUSY vacua. Thus, a compactification of 4d N = 2 theory T [C] on a circle yields a 3d N = 4 sigma-model whose target is the hyper-Kähler manifold MSU SY (T [C], G) = MH (G, C) (5.105) while a 3-manifold bounded by C defines a half-BPS boundary condition, i.e. a brane in the sigmamodel language. More precisely, a 3-manifold M3 with C = ∂M3 gives rise to a brane of type (A, B, A) with respect to the hyper-Kähler structure on MH (G, C). It is supported on a mid-dimensional submanifold of MH (G, C) which can be identified with the moduli space of flat GC connections on M3 : Mflat (M3 , GC ) ⊂ MH (G, C) (5.106) 80 Note, according to (5.5), the space of flat GC connections on M3 is precisely the space of SUSY vacua (parameters) of the 3d N = 2 theory T [M3 ] on a circle. When combined with (5.105) this gives MSUSY (T [M3 ], G) ⊂ MSUSY (T [C], G) (5.107) In this description, the mapping class group of the Riemann surface C (which we already identified with the duality group of T [C]) acts by autoequivalences on branes in the sigma-model with the target space MH (G, C). See [95, 11] for various examples of the mapping class group action on (A, B, A) branes in the Hitchin moduli space. In particular, when G = SU (2) and C = T 2 is a 2-torus, the Hitchin moduli space is a flat hyperKähler space MH (G, C) ∼ = (C∗ × C∗ )/Z2 parametrized by C∗ -valued holonomy eigenvalues x and y modulo the Weyl group action. This is also the space of vacua of T [C, G] after dimensional reduction on a circle. Each 3-manifold with a toral boundary defines a middle-dimensional submanifold or an (A, B, A) brane. Thus, when translated to language of geometry, the boundary conditions (5.99) and (5.100) correspond to (A, B, A) branes supported on x = 1 and y = 1, respectively: |x = 1i = |D5i (5.108) |y = 1i = |NS5i Similarly, the duality wall of type φ = S is a “correspondence” Mflat (M3 , GC ) ⊂ MH (G, C) × MH (G, C) associated with the mapping cylinder M3 ∼ = C × I, x+ 1 1 = y0 + 0 x y , y+ 1 1 = x0 + 0 y x (5.109) that exchanges the SL(2, C) holonomies on a- and b-cycles of C = T 2 . Note, these relations are deformed in MSUSY (T [M3 ], G) ⊂ MSUSY (T [C], G) × MSUSY (T [C], G) for a generic value of the fugacity t. 5.6.1.2 Lens space theories and matrix models In the above discussion we used the solid torus (5.97)–(5.98) as a simple example of a handlebody, in this case bounded by C = T 2 . Likewise, the simplest example of a closed 3-manifold obtained by gluing two solid tori is the Lens space L(p, 1) = h0|ST p S|0i ∼ = S 3 /Zp (5.110) 81 Using the dictionary (5.99) and (5.100), we can identify the corresponding 3d N = 2 theory T [L(p, 1)] as the theory on D3-branes suspended between a NS5-brane and a (p, 1)-fivebrane: T [L(p, 1); G] = SUSY G−p Chern-Simons theory (5.111) Following [58], here we assumed that the gauge group G is of Cartan type A, i.e. G = U (N ) or G = SU (N ). It would be interesting, however, to test the conjecture (5.111) for other groups G. Now, let us discuss this gluing more carefully, first from the viewpoint of flat connections (= SUSY vacua) and then from the viewpoint of partition functions. According to (5.100) and (5.108), the solid torus boundary condition S|0i in N = 4 super-Yang-Mills T [C] imposes a Neumann boundary condition on x and a Dirichlet boundary condition on y. In fact, the solid torus theory here is basically the theory of the unknot, T [01 ], discussed in section 5.4.2. Its supersymmetric parameter space (5.43) is a linear subspace of MSUSY (T [C], G) defined by y = 1. Note, the equation y − 1 = 0 is precisely the defining equation of the abelian branch, which in our present example is the entire moduli space Mflat (M3 , GC ) = MSUSY (T [M3 ], G). Therefore, had we ignored this component, the space of SUSY vacua would be completely empty, both for the solid torus theory T [S 1 × D2 ] and for everything else that can be obtained from it by gluing! A concatenation of the T p duality wall with this boundary condition adds a supersymmetric Chern-Simons term at level p for the global U (1)x symmetry of the theory T [01 ]. If we are only interested in SUSY vacua and parameters of a theory T [M3 ] (= flat connections on M3 ) we need to know how this operation affects the effective twisted superpotential, which for a general theory T [M3 ] has a simple form: Tp : f→W f + p (log x)2 W 2 (5.112) For the case at hand, the result of this operation modifies the space of SUSY parameters from y = 1 to y = xp . Finally, gluing h0|S and T p S|0i in (5.110) means sandwiching N = 4 super-Yang-Mills between the corresponding boundary conditions. In our IR description of the boundary conditions, this makes U (1)x dynamical, so that all critical points of the effective twisted superpotential [58]: fT [M ] = W f f W 3 T [M − ] − Wφ ◦ T [M + ] 3 3 (5.113) become SUSY vacua (= flat connections) of the theory T [M3 ] associated with the gluing (5.101). For the Lens space (5.110), we get a set of points {y = 1} ∪ {y = xp } in (C∗ × C∗ )/Z2 , which are in one-to-one correspondence with massive SUSY vacua of the Lens space theory T [L(p, 1); SU (2)].14 More generally, for G = U (N ) the flat connections on L(p, 1) or, equivalently, the SUSY vacua of (5.111) are labeled by Young diagrams ρ with at most p − 1 rows and N columns, i.e. Young 14 In section 5.6.1.3, we will take a look at a more complicated gluing operation where this simple analysis of vacua encounters some subtleties. 82 diagrams that fit in a rectangle of size N ×(p−1). Note, these are in one-to-one correspondence with b N (equivalently, of b integrable representations of su(p) u(N )p ), the fact that plays an important role [96, 97, 98, 99] in the study of Vafa-Witten partition function on ALE spaces bounded by L(p, 1). Next, let us consider the gluing (5.110) at the level of partition functions. The partition function of the theory (5.111) on the ellipsoid Sb3 is given by (see e.g. [100]): ZSb3 Z Y r Y 1 dσi e−iπpσ·σ 4 sinh(πbσ · α) sinh(πb−1 σ · α) |W| i=1 α∈Λ+ rt Y iπ iπ 2 πα · ρ −2 r/2 exp dim G − (b + b )h dim G p 2 sin 4 12p p + = = (5.114) α∈Λrt where r = rank(G), W is the Weyl group of G, h is the dual Coxeter number of G, Λ+ rt is the set of positive roots of G, and ρ is the Weyl vector (half the sum of the positive roots). Furthermore, turning on a FI parameter ζ contributes an extra term e4πiζTr σ into the integral (5.114). As usual, the Sb3 partition function of T [L(p, 1); G] should have the following structure ZSb3 = X ρ ||B ρ (q)||2S (5.115) 2 ρ where q = e~ = e2πib and each block B ρ (q) ∼ ZCS (L(p, 1); G) is expected to represent the Chern- Simons partition function computed in the background of a flat connection labeled by ρ. (Recall from our earlier discussion that classical solutions in Chern-Simons theory on M3 = L(p, 1) are labeled by certain Young diagrams ρ.) Unfortunately, there is no systematic algorithm to define holomorphic blocks in general 3d N = 2 theories and, as a result, the factorization (5.115) is not known at present for the N = 2 superChern-Simons theory (5.111). However, the integral form of the partition function (5.114) does share many key features with the Chern-Simons partition function on the Lens space that will be discussed in the next section (and extended to more general Seifert manifolds). Here, let us just note that log ZSb3 in (5.114) has the form of a power series in ~ = 2πib2 that starts with the leading ~1 term and terminates at the order O(~). This is indeed the property of Chern-Simons partition function on L(p, 1): according to a famous result of Lawrence and Rozansky [101], higher loop corrections to ρ ZCS (L(p, 1); G) all vanish. Finally, we propose a “lift” of the gluing formula (5.113) to a similar formula at the level of partition functions, cf. (5.102): Z ZT [M3 ] = [dU (x)] ZT [M − ] (x) · Zφ ◦ T [M + ] (x−1 ) 3 3 (5.116) where the integration measure [dU ] = ZT [C] dx is determined by the 4d N = 2 theory T [C; G] 83 associated with the Riemann surface C = ∂M3+ = −∂M3− . It would be interesting to test this gluing formula in concrete examples, including the Lens spaces and Seifert manifolds discussed here. Note that with the t-variable that keeps track of homological grading, (5.116) basically is a surgery formula for homological knot invariants. Such formulas are indeed known in the context of knot Floer homology and its version for general 3-manifolds, the Heegaard Floer homology. As explained around (5.101), we can construct closed 3-manifolds by gluing open 3-manifolds along their boundaries. The Chern-Simons partition functions on manifolds with torus boundary depend on a parameter x, which should be integrated out upon gluing. For a particular class of 3-manifolds, the resulting Chern-Simons partition functions can be represented as matrix integrals, much like (5.114), where the integration measure is responsible for integrating out the parameters x. The integrands of such matrix models take the form 1 exp − V (x) , ~ (5.117) 2πi where V (x) is usually called potential and 2πi ~ = log q is called the “level”. Let us note that in the case of 3-manifolds with boundary, when the parameters x are not integrated out, the same representation f + . . .) was used to read off the twisted superpotentials of dual of partition functions ZCS ∼ exp( ~1 W N = 2 theories, such as (5.46) or (5.51). One is therefore tempted to postulate, that a matrix model potential V (x) might encode information about the twisted superpotential and field content of the dual N = 2 theory T [M3 ] associated to a closed 3-manifold M3 . Let us demonstrate that this is indeed the case. For non-abelian Chern-Simons theories it is convenient to a write matrix model representation of their partition functions in terms of eigenvalues σi = log xi . A very well known example is a matrix model representation of the U (N ) Chern-Simons partition function on M3 = S 3 [76, 102], whose measure takes the form of a trigonometric deformation of the Vandermonde determinant, and the potential V (σ) = σ 2 /2 is Gaussian in σ = log x. More generally, the matrix model potential for M3 = L(p, 1) and G = U (N ) takes the form V (σ) = pσ 2 /2. More involved integral representations of Chern-Simons partition functions on other Lens spaces and Seifert homology spheres can be found in [101, 76, 102]. Various other matrix integral representations of Chern-Simons or related topological string partition functions, including the refined setting, were constructed in [103, 104, 105, 106, 107, 72, 108, 109]. Let us now consider more seriously the proposal that the potential of a Chern-Simons matrix model determines the dual 3d N = 2 theory T [M3 ]. For example, as reviewed above, the potential for a theory of the Lens space L(p, 1) takes the form V (σ) = pσ 2 /2. Taking into account a minus sign in (5.117), and using by now familiar 3d-3d dictionary, we might conclude that the dual theory is N = 2 theory at level −p, at least in the abelian case. Due to the universal form of the matrix 84 integral, we might also be tempted to declare that in the nonabelian case the dual theory is U (N ) theory at level −p. This is precisely the dual theory (5.111) which was originally constructed by other means. We also emphasize that the form of the matrix model reflects the structure of the gluing (5.101), namely the fact that the resulting Lens space (5.110) is constructed from two solid tori (unknot complements), glued with a suitable SL(2, Z) twist φ. Indeed, in this case the potential factor (5.117) represents the gluing SL(2, Z) element φ, while the information about two solid tori is encoded in the matrix model measure. This construction is discussed in detail e.g. in [102]. We can do similarly for Seifert manifold M3 at least G = U (1) [1]. 5.6.1.3 Dehn surgery As a final simple illustration of the necessity of accounting for all flat connections, we return to the basic Dehn surgery operation (5.103). Suppose that the knot K = 31 is the trefoil. As we know well from section 5.4.2, the A-polynomial15 for the trefoil, parametrizing MSUSY (T [31 ], SU (2)) for the full trefoil-complement theory T [31 , SU (2)], is A(x, y) = (y − 1)(y + x6 ) ⊂ (C∗ × C∗ )/Z2 . (5.118) Here x and y are the C∗ -valued eigenvalues of longitude and meridian SL(2, C) holonomies on the torus boundary of the knot complement, well defined up to the Weyl-group action (x, y) 7→ (x−1 , y −1 ). We recall that the (y − 1) component of the A-polynomial corresponds to an abelian flat connection on the knot complement, while the (y + x6 ) component corresponds to an irreducible flat connection. Suppose that we perform Dehn surgery with q/p = ±1 on the trefoil knot complement. The result is a Brieskorn sphere Σ[2, 3, 5] and Σ[2, 3, 7]; 3 Sp/q (31 ) =   Σ[2, 3, 5] p/q = +1 (5.119)  Σ[2, 3, 7] p/q = −1 The Brieskorn sphere is a closed 3-manifold and is defined as Σ[a, b, c] := {(x, y, z) ∈ C3 |xa + y b + z c = 0} ∩ S 5 (5.120) where a, b, and c are coprime. 3 In each case, the moduli space of flat SL(2, C) connections on Sp/q (31 ), consists of isolated 15 In contrast to the rest in this chapter, we take care in this section to write A-polynomials in terms of actual SL(2, C) meridian and longitude eigenvalues rather than their squares. Thus, for the trefoil, the non-abelian Apolynomial is written as y + x6 rather than y + x3 . The distinction is important for consistently counting SL(2, C) (as opposed to P SL(2, C), etc.) flat connections resulting from surgery. 85 points. It is easy to count them directly from a presentation of the fundamental groups of the Brieskorn spheres, π1 (Σ[2, 3, 5]) = ha, b | a3 = b5 = (ab)2 i , π1 (Σ[2, 3, 7]) = ha, b | a3 = b7 = (ab)2 i . (5.121) 3 3 We find |Mflat (S+1 (31 ), SL(2, C))| = 3 and |Mflat (S−1 (31 ), SL(2, C))| = 4. These counts must 3 equal the numbers of isolated vacua of the theories T [S±1 (31 ), SU (2)] on R2 × S 1 . Now compare the count of flat connections on the Brieskorn spheres with the intersection points of the varieties (xp y q = 1) ∩ (A(x, y) = 0) =   4 points p/q = 1  5 points p/q = −1 . (5.122) This does not quite match the count of flat connections on the Brieskorn spheres: in each case, there is one extra intersection point in (5.122). In particular, in each case, the intersection point (x, y) = (−1, −1) corresponds to flat connections on the knot complement S 3 \31 and the solid surgery torus whose eigenvalues match at the T 2 surgery interface, but whose full holonomies do not. Namely, the flat connection on the solid surgery torus with eigenvalues (−1, −1) is trivial, while the flat connection on the trefoil knot complement with eigenvalues (−1, −1) is parabolic, meaning 1 the full holonomy matrix is −1 0 −1 . This is not unexpected, since (x, y) = (−1, −1) lies on the nonabelian branch y +x6 = 0 of the trefoil’s A-polynomial. After subtracting the “false” intersection 3 (31 ) point from the counts in (5.122), we recover the expected number of flat connections on S+1 3 and S−1 (31 ). Physically, (5.122) is the (naively) expected count of vacua when gluing the trefoil theory to an unknot theory with the appropriate element φ ∈ SL(2, Z) corresponding to the Dehn surgery. The presence of a “false” intersection point (x, y) = (−1, −1) suggests that the corresponding vacuum in the glued theory must be lifted. It would be interesting to uncover the mechanism behind this. The remaining vacua match the count of flat connections on the Brieskorn spheres (i.e. vacua of 3 T [S±1 (31 ), SU (2)]), as they should. Crucially the vacuum corresponding to the intersection point (x, y) = (1, 1) must be included in order for the count to work out; this intersection point sits on 3 the abelian branch (y − 1) of the trefoil A-polynomial, and labels the trivial flat connection on S±1 . A similar phenomenon occurs when considering simple surgeries on the figure-eight knot complement S 3 \41 . For example, the Brieskorn sphere Σ[2, 3, 7] may be constructed from +1 or −1 surgeries on S 3 \41 . (The two different surgeries produce opposite orientations on Σ[2, 3, 7].) The in tersection of the full figure-eight A-polynomial A(x, y) = (y −1) x4 −(1−x2 −2x4 −x6 +x8 )y +x4 y 2 with the surgery conditions xy ±1 = 1 yield (xy ±1 = 1) ∩ (A(x, y) = 0) = 5 points . (5.123) 86 Four of these five intersection points, including the point on the abelian branch y −1 = 0, correspond to the expected flat SL(2, C) connections on Σ[2, 3, 7]. The fifth intersection point, at (x, y) = (−1, −1), does not correspond to any flat connection on Σ[2, 3, 7], because the connection with eigenvalues (x, y) = (−1, −1) on the knot complement is parabolic, while on the solid surgery torus it would have to be trivial. Explicitly, the meridian and longitude holonomies of the connections √ 1 −1 ±2i 3 , which will never on S 3 \41 with (x, y) = (−1, −1) are conjugate to µ = −1 0 −1 , λ = 0 −1 satisfy µp λq = I for any p, q. 5.6.2 Boundary conditions in 3d N = 2 theories So far we discussed what happens when 3-manifolds have boundaries, along which they can be glued, cf. (5.101). Now let us briefly discuss what happens when the space-time of 3d N = 2 theory T [M3 ; G] has a boundary. (0,2) multiplet contribution to half-index chiral θ(−q 2 x; q)−1 Fermi θ(−q 2 x; q) U (N ) gauge R−1 R (q; q)2N ∞ Q i6=j θ(−q − 21 σi /σj ; q) Table 5.1: Building blocks of 2d boundary theories and their contributions to the half-index. Much like in Chern-Simons theory on M3 the presence of non-trivial boundary requires specifying boundary conditions, the same is true in the case of 3d N = 2 theories. One important novelty, though, is that some boundary conditions are now distinguished if they preserve part of supersymmetry, such as half-BPS boundary conditions that preserve N = (0, 2) supersymmetry on the boundary. These “B-type” boundary conditions have been studied only recently in [62] and then in [110]. In the presence of a boundary (or, more generally, a domain wall) one can define a generalization of the index as a partition function on S 1 ×q D with a prescribed B-type boundary condition on the boundary torus S 1 ×q S 1 ∼ = T 2 of modulus τ , as illustrated in Figure 5.1. The resulting half-index IS 1 ×q D is essentially a convolution of the flavored elliptic genus of the 2d N = (0, 2) boundary theory with the index of a 3d N = 2 theory on S 1 ×q D. The contribution of (0, 2) boundary degrees of freedom is summarized in Table 5.1 where, as usual, gauge symmetries result in integrals over the corresponding variables σi . The half-index IS 1 ×q D labeled by a particular choice of the boundary condition can be viewed as a UV counterpart of a holomorphic block labeled by a choice of the massive vacuum in the IR. 87 Moreover, since the half-index is invariant under the RG flow, it makes sense to identify some of massive vacua and integration contours in the IR theory with specific boundary conditions in the UV. The latter, in turn, can sometimes be identified with 4-manifolds via (5.9), which altogether leads to an interesting correspondence between certain holomorphic blocks and 4-manifolds. Note, that for theories T [M3 ; G] labeled by closed 3-manifolds, supersymmetric vacua ρ ∈ MSUSY (T [M3 ; G]) specify boundary conditions for the Vafa-Witten topological gauge theory on a 4-manifold bounded by M3 . Therefore, had we missed any of the vacua in constructing T [M3 ; G] there would be no hope to relate supersymmetric boundary and 4-manifolds in (5.9). For instance, let us consider one of the simplest 3d N = 2 theories, namely the super-ChernSimons theory with gauge group G = U (N ) that in (5.111) we identified with the Lens space theory. As we mentioned earlier, the holomorphic blocks for this theory are not known. However, their UV counterparts IS 1 ×q D are easy to write down by choosing various B-type boundary conditions constructed in [62, 58, 110]. Thus, a simple boundary condition involves pN Fermi multiplets on the boundary. According to the rules in Table 5.1, its flavored elliptic genus can be interpreted as the half-index of 3d N = 2 super-Chern-Simons theory (5.111) with gauge group G = U (N ): pN IS 1 ×q D = q − 24 p Y N Y θ(xi zj ; q) (5.124) i=1 j=1 Moreover, it can be identified with the Vafa-Witten partition function of the ALE space −2 Ap−1 = M4 (su(p)) = M4 (−2 •− · · · −• ) | {z } (5.125) p−1 written in the “continuous basis” b u(N )p U (N ) (5.126) ZVW [Ap−1 ]ρ (q, x) = χsρcu(p)N (q, x) (5.127) U (N ) ZVW [Ap−1 ](q, x|z) := X χρt (q, z) ZVW [Ap−1 ]ρ (q, x) ρ where U (N ) is the well known form of the Vafa-Witten partition function on the ALE space (5.125) written in the “discrete basis” [96, 97, 98, 99]. Here, ρ is a Young diagram with at most p − 1 rows and N columns that in the previous section we identified with the choice of flat connection on M3 = ∂M4 = L(p, 1). 88 Bibliography [1] H.-J. Chung, T. Dimofte, S. Gukov, and P. Sulkowski, 3d-3d Correspondence Revisited, arXiv:1405.3663. [2] O. Aharony, A. Hanany, K. Intriligator, N. Seiberg, and M. J. Strassler, Aspects of N=2 Supersymmetric Gauge Theories in Three Dimensions, Nucl. Phys. B499 (1997), no. 1-2 67–99, [hep-th/9703110v1]. [3] J. de Boer, K. Hori, H. Ooguri, and Z. Yin, Mirror Symmetry in Three-Dimensional Gauge Theories, SL(2,Z) and D-Brane Moduli Spaces, Nucl. Phys. B493 (1996) 148–176, [hep-th/9612131v1]. [4] A. N. Redlich, Gauge noninvariance and parity nonconservation of three-dimensional fermions, Phys. Rev. Lett. 52 (1984), no. 1 18–21. [5] A. N. Redlich, Parity violation and gauge noninvariance of the effective gauge field action in three dimensions, Phys. Rev. D 29 (1984), no. 10 2366–2374. [6] L. Alvarez-Gaumé and E. Witten, Gravitational anomalies, Nucl. Phys. B234 (1984), no. 2 269–330. [7] V. Borokhov, A. Kapustin, and X.-k. Wu, Topological disorder operators in three-dimensional conformal field theory, JHEP 0211 (2002) 049, [hep-th/0206054]. [8] V. Borokhov, A. Kapustin, and X.-k. Wu, Monopole operators and mirror symmetry in three-dimensions, JHEP 0212 (2002) 044, [hep-th/0207074]. [9] K. Intriligator and N. Seiberg, Aspects of 3d N=2 Chern-Simons-Matter Theories, JHEP 1307 (2013) 079, [arXiv:1305.1633]. [10] N. A. Nekrasov and S. L. Shatashvili, Supersymmetric vacua and Bethe ansatz, Nucl. Phys. B, Proc. Suppl. 192-193 (2009) 91–112, [arXiv:0901.4744]. [11] T. Dimofte and S. Gukov, Chern-Simons Theory and S-duality, arXiv:1106.4550. 89 [12] E. Witten, Phases of N=2 Theories In Two Dimensions, Nucl. Phys. B403 (1993) 159–222, [hep-th/9301042v3]. [13] C. Beem, T. Dimofte, and S. Pasquetti, Holomorphic Blocks in Three Dimensions, arXiv:1211.1986. [14] V. Pestun, Localization of gauge theory on a four-sphere and supersymmetric Wilson loops, arXiv:0712.2824. [15] A. Kapustin, B. Willett, and I. Yaakov, Exact Results for Wilson Loops in Superconformal Chern-Simons Theories with Matter, JHEP 1003 (2010) 089, [arXiv:0909.4559]. Published in: JHEP 1003:089,2010 32 pages. [16] D. L. Jafferis, The Exact Superconformal R-Symmetry Extremizes Z, arXiv:1012.3210. [17] N. Hama, K. Hosomichi, and S. Lee, Notes on SUSY Gauge Theories on Three-Sphere, JHEP 1103 (2011) 127, [arXiv:1012.3512]. [18] N. Hama, K. Hosomichi, and S. Lee, SUSY Gauge Theories on Squashed Three-Spheres, arXiv:1102.4716. [19] S. Kim, The complete superconformal index for N=6 Chern-Simons theory, Nucl. Phys. B821 (2009) 241–284, [arXiv:0903.4172]. [20] Y. Imamura and S. Yokoyama, Index for three dimensional superconformal field theories with general R-charge assignments, arXiv:1101.0557. [21] A. Kapustin and B. Willett, Generalized Superconformal Index for Three Dimensional Field Theories, arXiv:1106.2484. [22] D. L. Jafferis, I. R. Klebanov, S. S. Pufu, and B. R. Safdi, Towards the F-Theorem: N=2 Field Theories on the Three-Sphere, JHEP 1106 (2011) 102, [arXiv:1103.1181]. [23] C. Closset, T. T. Dumitrescu, G. Festuccia, Z. Komargodski, and N. Seiberg, Contact Terms, Unitarity, and F-Maximization in Three-Dimensional Superconformal Theories, arXiv:1205.4142. [24] S. Pasquetti, Factorisation of N = 2 theories on the squashed 3-sphere, arXiv:1111.6905. [25] T. Dimofte, D. Gaiotto, and S. Gukov, 3-Manifolds and 3d Indices, arXiv:1112.5179. [26] S. Cecotti and C. Vafa, Topological-anti-topological fusion, Nucl. Phys. B367 (1991), no. 2 359–461. 90 [27] S. Gukov, Three-Dimensional Quantum Gravity, Chern-Simons Theory, and the A-Polynomial, Commun. Math. Phys. 255 (2005), no. 3 577–627, [hep-th/0306165v1]. [28] T. Dimofte, S. Gukov, J. Lenells, and D. Zagier, Exact Results for Perturbative Chern-Simons Theory with Complex Gauge Group, Comm. Num. Thy. and Phys. 3 (2009), no. 2 363–443, [arXiv:0903.2472]. [29] T. Dimofte, Quantum Riemann Surfaces in Chern-Simons Theory, arXiv:1102.4847. [30] E. Witten, Analytic Continuation of Chern-Simons Theory, arXiv:1001.2933. [31] E. Witten, Quantum Field Theory and the Jones Polynomial, Comm. Math. Phys. 121 (1989), no. 3 351–399. [32] S. Elitzur, G. Moore, A. Schwimmer, and N. Seiberg, Remarks on the Canonical Quantization of the Chern-Simons-Witten Theory, Nucl. Phys. B326 (1989), no. 1 108–134. [33] E. Witten, Quantization of Chern-Simons Gauge Theory with Complex Gauge Group, Comm. Math. Phys 137 (1991) 29–66. [34] D. Cooper, M. Culler, H. Gillet, D. Long, and P. Shalen, Plane Curves Associated to Character Varieties of 3-Manifolds, Invent. Math. 118 (1994), no. 1 47–84. [35] T. D. Dimofte, Refined BPS invariants, Chern-Simons theory, and the quantum dilogarithm. PhD thesis, California Institute of Technology, 2010. [36] S. Gukov and P. Sulkowski, A-polynomial, B-model, and Quantization, arXiv:1108.0002. [37] H. Fuji, S. Gukov, and P. Sulkowski, Volume Conjecture: Refined and Categorified, arXiv:1203.2182. [38] H. Fuji, S. Gukov, and P. Sulkowski, Super-A-polynomial for knots and BPS states, arXiv:1205.1515. [39] T. Dimofte, D. Gaiotto, and S. Gukov, Gauge Theories Labelled by Three-Manifolds, arXiv:1108.4389. [40] W. D. Neumann and D. Zagier, Volumes of hyperbolic three-manifolds, Topology 24 (1985), no. 3 307–332. [41] L. D. Faddeev, Discrete Heisenberg-Weyl Group and Modular Group, Lett. Math. Phys. 34 (1995), no. 3 249–254. [42] D. Shale, Linear Symmetries of Free Boson Fields, Trans. Amer. Math. Soc. 103 (1962) 149–167. 91 [43] A. Weil, Sur Certains Groupes d’Opérateurs Unitaires, Acta Math. 111 (1964) 143–211. [44] L. F. Alday, D. Gaiotto, and Y. Tachikawa, Liouville Correlation Functions from Four-Dimensional Gauge Theories, Lett. Math. Phys. 91 (2010), no. 2 167–197, [arXiv:0906.3219]. [45] T. Dimofte, M. Gabella, and A. B. Goncharov, K-Decompositions and 3d Gauge Theories, arXiv:1301.0192. [46] S. Cecotti, C. Cordova, and C. Vafa, Braids, Walls, and Mirrors, arXiv:1110.2115. [47] C. Cordova and D. L. Jafferis, Complex Chern-Simons from M5-branes on the Squashed Three-Sphere, arXiv:1305.2891. [48] S. Lee and M. Yamazaki, 3d Chern-Simons Theory from M5-branes, arXiv:1305.2429. [49] N. Drukker, D. Gaiotto, and J. Gomis, The Virtue of Defects in 4D Gauge Theories and 2D CFTs, arXiv:1003.1112. [50] K. Hosomichi, S. Lee, and J. Park, AGT on the S-duality Wall, JHEP 1012 (2010) 079, [arXiv:1009.0340]. [51] Y. Terashima and M. Yamazaki, SL(2,R) Chern-Simons, Liouville, and Gauge Theory on Duality Walls, arXiv:1103.5748. [52] T. Dimofte, S. Gukov, and L. Hollands, Vortex Counting and Lagrangian 3-manifolds, arXiv:1006.0977. [53] H. Ooguri and C. Vafa, Knot Invariants and Topological Strings, Nucl. Phys. B5777 (Jan, 2000) 419–438, [hep-th/9912123v3]. [54] R. Gopakumar and C. Vafa, M-Theory and Topological Strings-I, hep-th/9809187v1. [55] M. Aganagic and C. Vafa, Large N Duality, Mirror Symmetry, and a Q-deformed A-polynomial for Knots, arXiv:1204.4709. [56] H. Fuji, S. Gukov, M. Stosic, and P. Sulkowski, 3d analogs of Argyres-Douglas theories and knot homologies, arXiv:1209.1416. [57] S. Nawata, P. Ramadevi, Zodinmawia, and X. Sun, Super-A-polynomials for Twist Knots, JHEP 1211 (2012) 157, [arXiv:1209.1409]. [58] A. Gadde, S. Gukov, and P. Putrov, Fivebranes and 4-manifolds, arXiv:1306.4320. [59] D. Gaiotto, L. Rastelli, and S. S. Razamat, Bootstrapping the superconformal index with surface defects, arXiv:1207.3577. 92 [60] M. Bullimore, M. Fluder, L. Hollands, and P. Richmond, The superconformal index and an elliptic algebra of surface defects, arXiv:1401.3379. [61] S. S. Razamat and B. Willett, Down the rabbit hole with theories of class S, arXiv:1403.6107. [62] A. Gadde, S. Gukov, and P. Putrov, Walls, Lines, and Spectral Dualities in 3d Gauge Theories, arXiv:1302.0015. [63] M. Khovanov, A categorification of the Jones polynomial, Duke Math. J. 101 (2000) 359426. [64] M. Khovanov and L. Rozansky, Matrix factorizations and link homology, math/0401268. [65] Y. Yonezawa, Quantum (sln , ∧Vn ) link invariant and matrix factorizations, Nagoya Math. J. 204 (2011) 69–123, [arXiv:0906.0220]. [66] H. Wu, A colored sl(N)-homology for links in S 3 , arXiv:0907.0695. [67] N. M. Dunfield, S. Gukov, and J. Rasmussen, The Superpolynomial for knot homologies, math/0505662. [68] M. Khovanov and L. Rozansky, Matrix factorizations and link homology II, math/0505056. [69] S. Gukov and M. Stosic, Homological algebra of knots and BPS states, arXiv:1112.0030. [70] P. Freyd, D. Yetter, J. Hoste, W. B. R. Lickorish, K. Millett, and A. A. Ocneanu, A new polynomial invariant of knots and links, Bull. Am. Math. Soc. 12 (1985) 239–246. [71] P. Dunin-Barkowski, A. Mironov, A. Morozov, A. Sleptsov, and A. Smirnov, Superpolynomials for toric knots from evolution induced by cut-and-join operators, arXiv:1106.4305. [72] M. Aganagic and S. Shakirov, Knot Homology from Refined Chern-Simons Theory, arXiv:1105.5117. [73] I. Cherednik, Jones polynomials of torus knots via DAHA, arXiv:1111.6195. [74] S. Gukov, A. Iqbal, C. Kozcaz, and C. Vafa, Link Homologies and the Refined Topological Vertex, arXiv:0705.1368. [75] H. Itoyama, A. Mironov, A. Morozov, and A. Morozov, HOMFLY and superpolynomials for figure eight knot in all symmetric and antisymmetric representations, arXiv:1203.5978. [76] M. Marino, Chern-Simons theory, matrix integrals, and perturbative three manifold invariants, Commun.Math.Phys. 253 (2004) 25–49, [hep-th/0207096]. 93 [77] S. de Haro, Chern-Simons theory in lens spaces from 2-D Yang-Mills on the cylinder, JHEP 0408 (2004) 041, [hep-th/0407139]. [78] C. Beasley and E. Witten, Non-Abelian localization for Chern-Simons theory, J.Diff.Geom. 70 (2005) 183–323, [hep-th/0503126]. [79] S. de Haro, A Note on knot invariants and q-deformed 2-D Yang-Mills, Phys.Lett. B634 (2006) 78–83, [hep-th/0509167]. [80] M. Blau and G. Thompson, Chern-Simons Theory on Seifert 3-Manifolds, JHEP 1309 (2013) 033, [arXiv:1306.3381]. [81] S. Garoufalidis, On the Characteristic and Deformation Varieties of a Knot, Geom. Topol. Monogr. 7 (2004) 291–304, [math/0306230v4]. [82] M. Aganagic and C. Vafa, Large N Duality, Mirror Symmetry, and a Q-deformed A-polynomial for Knots, arXiv:1204.4709. [83] L. Ng, Framed knot contact homology, Duke Math. J. 141 (2008) 365–406, [math/0407071]. [84] L. Ng, Combinatorial knot contact homology and transverse knots, Adv. Math. 227 (2011) 2189–2219, [arXiv:1010.0451]. [85] H. Fuji and P. Sulkowski, Super-A-polynomial, arXiv:1303.3709. [86] S. Gukov and I. Saberi, Lectures on Knot Homology and Quantum Curves, arXiv:1211.6075. [87] Y. Terashima and M. Yamazaki, Semiclassical Analysis of the 3d/3d Relation, arXiv:1106.3066. [88] C. Cordova, S. Espahbodi, B. Haghighat, A. Rastogi, and C. Vafa, Tangles, Generalized Reidemeister Moves, and Three-Dimensional Mirror Symmetry, arXiv:1211.3730. [89] K. Intriligator and N. Seiberg, Mirror Symmetry in Three Dimensional Gauge Theories, Phys. Lett. B387 (1996) 513–519, [hep-th/9607207v1]. [90] D. Gaiotto and E. Witten, Knot Invariants from Four-Dimensional Gauge Theory, arXiv:1106.4789. [91] E. Gorsky, S. Gukov, and M. Stosic, Quadruply-graded colored homology of knots, arXiv:1304.3481. [92] T. Dimofte, D. Gaiotto, and R. van der Veen, RG Domain Walls and Hybrid Triangulations, arXiv:1304.6721. 94 [93] D.-E. Diaconescu, D-branes, monopoles and Nahm equations, Nucl.Phys. B503 (1997) 220–238, [hep-th/9608163]. [94] D. Gaiotto and E. Witten, Supersymmetric Boundary Conditions in N=4 Super Yang-Mills Theory, J. Stat. Phys. 135 (Dec, 2009) 789–855, [arXiv:0804.2902]. [95] S. Gukov, Gauge theory and knot homologies, Fortsch.Phys. 55 (2007) 473–490, [arXiv:0706.2369]. [96] H. Nakajima, Instantons on ALE spaces, quiver varieties, and Kac-Moody Algebras, Duke Math. 76 (1994) 365–416. [97] C. Vafa and E. Witten, A Strong coupling test of S duality, Nucl.Phys. B431 (1994) 3–77, [hep-th/9408074]. [98] R. Dijkgraaf, L. Hollands, P. Sulkowski, and C. Vafa, Supersymmetric Gauge Theories, Intersecting Branes and Free Fermions, JHEP 0802 (Jan, 2008) 106, [arXiv:0709.4446]. [99] R. Dijkgraaf and P. Sulkowski, Instantons on ALE spaces and orbifold partitions, JHEP 0803 (2008) 013, [arXiv:0712.1427]. [100] O. Aharony, S. S. Razamat, N. Seiberg, and B. Willett, 3d dualities from 4d dualities, JHEP 1307 (2013) 149, [arXiv:1305.3924]. [101] R. Lawrence and L. Rozansky, Witten-Reshetikhin-Turaev invariants of Seifert manifolds, Comm. Math. Phys. 205 (1999), no. 2 287–314. [102] M. Aganagic, A. Klemm, M. Marino, and C. Vafa, Matrix model as a mirror of Chern-Simons theory, JHEP 0402 (2004) 010, [hep-th/0211098]. [103] S. de Haro and M. Tierz, Discrete and oscillatory matrix models in Chern-Simons theory, Nucl.Phys. B731 (2005) 225–241, [hep-th/0501123]. [104] A. Klemm and P. Sulkowski, Seiberg-Witten theory and matrix models, Nucl.Phys. B819 (2009) 400–430, [arXiv:0810.4944]. [105] B. Eynard, A.-K. Kashani-Poor, and O. Marchal, A Matrix Model for the Topological String I: Deriving the Matrix model, arXiv:1003.1737. [106] H. Ooguri, P. Sulkowski, and M. Yamazaki, Wall Crossing As Seen By Matrix Models, Commun.Math.Phys. 307 (2011) 429–462, [arXiv:1005.1293]. [107] P. Sulkowski, Refined matrix models from BPS counting, Phys.Rev. D83 (2011) 085021, [arXiv:1012.3228]. 95 [108] R. J. Szabo and M. Tierz, q-deformations of two-dimensional Yang-Mills theory: Classification, categorification and refinement, Nucl.Phys. B876 (2013) 234–308, [arXiv:1305.1580]. [109] Z. Kknyesi, A. Sinkovics, and R. J. Szabo, Refined Chern-Simons theory and (q, t)-deformed Yang-Mills theory: Semi-classical expansion and planar limit, JHEP 1310 (2013) 067, [arXiv:1306.1707]. [110] T. Okazaki and S. Yamaguchi, Supersymmetric Boundary Conditions in Three Dimensional N = 2 Theories, Phys.Rev. D87 (2013) 125005, [arXiv:1302.6593].