About this Journal Submit a Manuscript Table of Contents
Advances in High Energy Physics
Volume 2011 (2011), Article ID 357016, 52 pages
http://dx.doi.org/10.1155/2011/357016
Review Article

BPS States, Crystals, and Matrices

1California Institute of Technology, Pasadena, CA 91125, USA
2Faculty of Physics, University of Warsaw, ul. Hoża 69, 00-681 Warsaw, Poland

Received 14 May 2011; Accepted 19 June 2011

Academic Editor: Amihay Hanany

Copyright © 2011 Piotr Sułkowski. This is an open access article distributed under the Creative Commons Attribution License, which permits unrestricted use, distribution, and reproduction in any medium, provided the original work is properly cited.

Abstract

We review free fermion, melting crystal, and matrix model representations of wall-crossing phenomena on local, toric Calabi-Yau manifolds. We consider both unrefined and refined BPS counting of closed BPS states involving D2- and D0-branes bound to a D6-brane, as well as open BPS states involving open D2-branes ending on an additional D4-brane. Appropriate limit of these constructions provides, among the others, matrix model representation of refined and unrefined topological string amplitudes.

1. Introduction

This paper is devoted to some aspects of counting of BPS states in a system of D𝑝-branes, with even 𝑝, in type IIA string compactifications. The problems of BPS counting span a vast area of research in supersymmetric gauge and string theories. Their important feature is a special, nonconstant character of BPS multiplicities: their values depend on various moduli and jump discontinuously along some special loci in the corresponding moduli space, so called walls of marginal stability. The pattern of these jumps follows wall-crossing formulas, found from physical perspective by Denef and Moore [1] and, in more general context, formulated mathematically by Kontsevich and Soibelman [2]. The regions of the moduli space in between walls of marginal stability, in which BPS multiplicities are (locally) constant, are called chambers.

The BPS states we are interested in, and which we will refer to as closed BPS states, arise as bound states of a single D6-brane with arbitrary number of D0 and D2-branes wrapping cycles of a toric Calabi-Yau space. More generally, we will also consider open BPS states, which arise when an additional D4-brane spans a Lagrangian submanifold inside the Calabi-Yau space and supports open D2-branes attached to it. The closed and open BPS states give rise, respectively, to single-particle states in the effective four-dimensional and two-dimensional theory (in remaining, space-time filling directions of, resp., D6 and D4-branes). In this context, the character of BPS multiplicities can be understood in much detail, and it relates to other interesting exactly solvable models: free fermions, crystal, and matrix models. In brief, these connections arise as follows. Firstly, BPS states we consider turn out to be in one-to-one correspondence with configurations of certain statistical models of melting crystals. The structure of these crystals depends on geometry of the underlying Calabi-Yau space, as well as on the chamber one is considering. In consequence, BPS counting functions, upon appropriate identification of parameters, coincide with generating functions of melting crystals. It turns out that the structure of these crystals can be given a free fermion representation. Furthermore, once such free fermion formulation is known, it can also be represented in terms of matrix models. Connection with vast theory of matrix models has many interesting mathematical and physical consequences and allows to shed new light on wall-crossing phenomena. The aim of this paper is to explain these connections.

The BPS generating functions which we consider are intimately related to topological string amplitudes on corresponding Calabi-Yau spaces. This relation is most transparent in the physical derivation discussed in Section 2, which relies on lifting the D-brane system to M-theory. The M-theory viewpoint makes contact with original formulation of closed topological strings by Gopakumar and Vafa [3, 4], and open topological strings by Ooguri and Vafa [5]. In particular, in one specific, so-called noncommutative chamber, the BPS-generating function is given as the modulus square of the topological string partition function. In all other chambers, BPS generating functions can be uniquely determined from that noncommutative result. There is also another special, so-called commutative chamber, in which BPS generating function coincides (up to the factor of MacMahon function) with the topological string partition function. For toric manifolds which we consider, such topological string amplitudes can be constructed, among the other, by means of the powerful topological vertex formalism [6]. Relation to crystal models was in fact first understood in this topological string chamber [79]. One advantage of the formalism presented in this paper is the fact that it allows to construct matrix model representation of all these generating functions (so, in particular, matrix model representation of topological string amplitudes).

In more detail, we will consider generating functions of D2 and D0-branes bound to a single D6-brane of the following form: 𝒵BPS𝑞𝑠=,𝑄𝛼,𝛽Ω(𝛼,𝛽)𝑞𝛼𝑠𝑄𝛽,(1.1) where 𝛼 is D0-brane charge and 𝛽𝐻2(𝑋,) is D2-brane charge. Multiplicities Ω(𝛼,𝛽) jump when central charges (which itself are functions of Kähler moduli) of building blocks of a bound state align, and therefore these generating functions are locally constant functions of Kähler moduli. Along the walls of marginal stability, the degeneracies Ω(𝛼,𝛽) change and indeed obey wall-crossing formulas of [1, 2] mentioned above.

If there is an additional D4-brane which spans a Lagrangian submanifold inside the Calabi-Yau space, in addition to the above closed BPS states, one can consider also open BPS states of D2-branes with boundaries ending on a one-cycle 𝛾 on this D4-brane. In this case, the BPS states arise on the remaining two-dimensional world-volume of the D4-brane. The holonomy of the gauge field along 𝛾 provides another generating parameter 𝑧, so that open BPS-generating functions take form 𝒵openBPS𝑞𝑠=,𝑄𝛼,𝛽,𝛾Ω(𝛼,𝛽)𝑞𝛼𝑠𝑄𝛽𝑧𝛾.(1.2) As we will show, generating functions of such open BPS states can be identified with integrands of matrix models mentioned above.

One more important aspect of BPS counting is referred to as refinement, and amounts to refining BPS counting by introducing one more parameter, customarily denoted 𝛽. The refinement can be introduced from several perspectives which give rise to identical results; however, their fundamental common origin is still not fully understood. We will introduce refinement by distinct counting of states with different 𝑆𝑈(2) spins inside spacetime 𝑆𝑂(4) rotation group in the generating function (1.1). In [10], it was argued that this physical viewpoint should agree with the mathematical counterpart of motivic deformation [2], and also a refined version of a crystal model was constructed. Another notion of refinement arises in Nekrasov partition functions, which are defined in a nontrivial gravitational (so-called Ω-) background parametrized by two parameters 𝜖1 and 𝜖2 [11]. Nekrasov partition functions can also be defined for five-dimensional gauge theories and then they agree with topological string amplitudes. In particular, the formalism of the topological vertex [6] has also been extended to the refined context in [12], and shown to reproduce relevant Nekrasov partition functions. Also BPS generating functions, in the limit of commutative chamber, are known to reproduce refined topological string amplitudes with 𝛽=𝜖1/𝜖2 [13]. However, the worldsheet definition of refined topological string amplitudes is not fully understood.

As an exemplary and, hopefully, inspiring application of the entire formalism presented in this paper, in the final Section 6 we derive matrix model representation of the refined topological string partition function for the conifold. The refined matrix model which we find has a standard measure; however, its potential is deformed by 𝛽-dependent terms. It is obtained by constructing appropriate refined crystal model and free fermion representation, and subsequently reformulating this representation in matrix model form. Finally, taking the limit of the commutative chamber, we obtain matrix model representation of the refined topological string amplitude. Even though we demonstrate this result in the conifold case, with some technical effort it can be generalized to other toric manifolds which we consider (As we recall in Section 6, refined topological string amplitudes were also postulated to be reproduced by another type of matrix models, so-called 𝛽-deformed ones (whose Vandermonde measure is deformed by raising it to power 𝛽); however, explicit computations showed that this cannot be the correct representation of refined amplitudes.).

1.1. Short Literature Guide

The literature on the topics presented in this paper is extensive and still growing, and we unavoidably mention just a fraction of important developments. The relation between Donaldson-Thomas invariants for the noncommutative chamber of the conifold was first found by Szendrői [14]. It was generalized to orbifolds of 3, and related to free fermion formalism, by Bryan and Young [15]. The relation to free fermions and crystals was extended to a large class of toric manifolds without compact four-cycles [16, 17]. These developments were accompanied by other mathematical works [18, 19].

In parallel to the above-mentioned mathematical activity, wall-crossing phenomena for local Calabi-Yau manifolds were analyzed from physical viewpoint. The analysis of nontrivial BPS counting for the conifold was described by Jafferis and Moore in [20]. This and more general cases were related to quivers and crystal models in [21, 22]. Derivation of BPS degeneracies from M-theory viewpoint and relation to closed topological strings were discussed in [23], and generalized to open BPS counting in [2427]. Relations to matrix models, discussed for plane partitions with some other motivation in [28], were extended to other crystal models relevant for BPS counting in [29], and also in [30]. Subsequently, it was related to open BPS counting in [27]. Refined BPS counting was related to crystal models in [10, 13], and corresponding matrix models were constructed in [31].

Let us also mention some other, related works devoted to crystals and free fermions. The fermionic construction of MacMahon function for 3 was originally presented in [7], and its relation to open topological strings and more complicated Calabi-Yau manifolds were discussed in [3234]. Newer ideas, analyzing more complicated systems involving D4-branes, were presented in [35, 36]. More expository presentations of various aspects described here can be found in [37, 38]. A general introduction to mathematical and physical aspects of mirror symmetry can be found in [39].

1.2. Plan

The plan of this paper is as follows. In Section 2, we introduce BPS generating functions and present one possible derivation of their form, which relies on the M-theory interpretation of a D-brane system, following [2325, 27]. In Section 3, we provide a little mathematical background and introduce notation pertaining to toric Calabi-Yau manifolds, free fermion formalism, and matrix models. In Section 4, we introduce fermionic formalism for BPS generating functions and present corresponding crystal models, building on earlier ideas of [7, 15] and following [16]. In Section 5, we reformulate the problem of closed BPS counting in terms of matrix models and relate it to open BPS counting [27, 29]. In Section 6, we refine our analysis, present refined BPS generating functions and crystals [10], and construct corresponding refined matrix models [31].

2. BPS Generating Functions

In this section, we introduce generating functions of BPS states of D-branes in toric Calabi-Yau manifolds. Our task in the rest of this paper is to provide interpretation of these generating functions in terms of free fermions, melting crystals, and matrix models. These generating functions can be derived using wall-crossing formulas, as was done first in the unrefined [20] and refined [10] conifold case, and later generalized to arbitrary geometry without compact four-cycles in [18, 19]. On the other hand, we will focus on a simpler physical derivation of BPS generating functions which uses the lift of the D-brane system to M-theory [23]. This also makes contact with M-theory interpretation of topological string theory and allows to express BPS counting functions in terms of topological string amplitudes. Moreover, this M-theory derivation can be extended to the counting of open BPS states,that is, open D2-branes attached to additional D4-brane, which we are also interested in [24, 25, 27].

We start this section by reviewing the M-theory derivation of (unrefined) closed and open BPS generating functions. Then, to get acquainted with a crystal interpretation of these generating functions, we discuss their crystal interpretation in simple cases of 3 and conifold. Later, using fermionic interpretation, we will generalize this crystal representation to a large class of toric manifolds without compact four cycles.

2.1. M-Theory Derivation

We start by considering a system of D2 and D0-branes bound to a single D6-brane in type IIA string theory. It can be reinterpreted in M-theory as follows [23]. When additional 𝑆1 is introduced as the eleventh dimension transversely to the D6-brane, then this D6-brane transforms into a geometric background of a Taub-NUT space with unit charge [40]. The Taub-NUT space is a circle fibration over 3, with a circle 𝑆1TN attaining a fixed radius 𝑅 at infinity, and shrinking to a point in the location of the original D6-brane. From M-theory perspective, bound states involving D2 and D0-branes are interpreted as M2-branes with momentum on a circle. Therefore, the counting of original bound states to the D6-brane is reinterpreted as the counting of BPS states of M2-branes in the Taub-NUT space. While in general this is still a nontrivial problem, for the purpose of counting BPS degeneracies we can take advantage of their invariance under continuous deformations of the Taub-NUT space, in particular under deformations of the radius 𝑅. We can therefore consider taking this radius to infinity, whereupon BPS counting is reinterpreted in terms of a gas of particles in 5. To make the problem fully tractable, we have to ensure that the particles are noninteracting, which would be the case if moduli of the Calabi-Yau would be tuned so that M2-branes wrapped in various ways would have aligned central charges. This can be achieved when Kähler parameters of the Calabi-Yau space are tuned to zero. However, to avoid generation of massless states, at the same time one has to include nontrivial fluxes of the M-theory three-form field through the two cycles of the Calabi-Yau and 𝑆1TN. In type IIA, this results in the 𝐵-field flux 𝐵 through two cycles of Calabi-Yau. Finally, to avoid creation of the string states arising from M5-branes wrapping four cycles in Calabi-Yau, we simply restrict considerations to manifolds without compact four cycles. For a state arising from D2-brane wrapping a class 𝛽, the central charge then reads 1𝑍(𝑙,𝛽)=𝑅(𝑙+𝐵𝛽),(2.1) where 𝑙 counts the D0-brane charge, which is taken positive to preserve the same supersymmetry.

Under the above conditions, the counting of D6-D2-D0 bound states is reinterpreted in terms of a gas of particles arising from M2-branes wrapped on cycles 𝛽. The excitations of these particles in 4, parametrized by two complex variables 𝑧1,𝑧2, are accounted for by the modes of the holomorphic field Φ𝑧1,𝑧2=𝑙1,𝑙2𝛼𝑙1,𝑙2𝑧𝑙11𝑧𝑙22.(2.2) Decomposing the isometry group of 4 as 𝑆𝑂(4)=𝑆𝑈(2)×𝑆𝑈(2), there are 𝑁𝛽𝑚,𝑚 five-dimensional BPS states of intrinsic spin (𝑚,𝑚). We are interested in their net number arising from tracing over 𝑆𝑈(2) spins 𝑁𝑚𝛽=𝑚(1)𝑚𝑁𝑚,𝑚𝛽.(2.3) The total angular momentum of a given state contributing to the index is 𝑙=𝑙1+𝑙2+𝑚. Finally, in a chamber specified by the moduli 𝑅 and 𝐵, the invariant degeneracies can be expressed as the trace over the corresponding Fock space 𝒵BPS=TrFock𝑞𝑄0𝑠𝑄𝑄2|||chamber=𝛽,𝑚𝑙1+𝑙2=𝑙1𝑞𝑙1+𝑙2𝑠+𝑚𝑄𝛽𝑁𝑚𝛽||||chamber=𝛽,𝑚𝑙=11𝑞𝑠𝑙+𝑚𝑄𝛽𝑙𝑁𝑚𝛽|||chamber,(2.4) where the subscript chamber denotes restriction to those factors in the above product, which represent states which are mutually BPS 𝑍(𝑙,𝛽)>0𝑞𝑠𝑙+𝑚𝑄𝛽<1.(2.5) As usual, 𝑄=𝑒𝑇 and 𝑞𝑠=𝑒𝑔𝑠 above encode, respectively, the Kähler class 𝑇 and the string coupling 𝑔𝑠 (we wish to distinguish carefully 𝑞𝑠 which encodes string coupling, from a counting parameter 𝑞 which will arise in what follows in crystal interpretation). The above condition on central charges is crucial in determining a particular form of the BPS generating functions. If we would restrict products in the formula (2.4) to factors with only positive 𝛽, we would get (up to possibly some factor of MacMahon function) the Gopakumar-Vafa representation of the topological string amplitude. With all negative and positive values of 𝛽, we would get modulus square of the topological string partition function. Therefore, the upshot of [23] is that in general the above BPS generating function can be expressed in terms of the closed topological string partition function 𝒵BPS=𝒵top(𝑄)𝒵top𝑄1||chamber,(2.6) where chamber restriction is to be understood as picking up only those factors in Gopakumar-Vafa product representation of 𝒵top for which (2.5) is satisfied. In this context, we will often refer to the choice of a chamber as a closed BPS chamber. The (instanton part of the) closed topological string partition function entering the above expression is given by [3, 4] 𝒵top𝑞(𝑄)=𝑀𝑠𝜒/2𝑙=1𝛽>0,𝑚1𝑄𝛽𝑞𝑠𝑚+𝑙𝑙𝑁𝑚𝛽,(2.7) where 𝑀(𝑞𝑠)=𝑙(1𝑞𝑙𝑠)𝑙 is the MacMahon function and 𝜒 is the Euler characteristic of the Calabi-Yau manifold.

To be more precise, an identification as a topological string partition function or its square arises if 𝑅>0 in (2.1). Because 𝑅 arises just as a multiplicative factor in (2.1), degeneracies depend only on its sign. Therefore, another extreme case corresponds to negative 𝑅 and 𝐵 sufficiently small, when only a single D6-brane contributes to the partition function 𝒵(𝑅<0,0<𝐵1)=1.(2.8) More generally, for 𝑅<0, BPS generating functions often (but not always) take finite form.

In what follows we denote BPS generating functions in chambers with positive 𝑅 by 𝒵, and in chambers with negative 𝑅 by 𝒵 (and often omit the subscript BPS). Topological string partition functions will be denoted by 𝒵top, while generating functions of melting crystals by ordinary 𝑍.

The above structure can be generalized by including in the initial D6-D2-D0 configuration additional D4-branes wrapping Lagrangian cycles in the internal Calabi-Yau manifold and extending in two space-time dimensions [24, 25, 27]. For simplicity, we consider a system with a single D4-brane wrapping a Lagrangian cycle. There are now additional BPS states in two remaining spacetime dimensions arising from open D2-branes ending on these D4-branes. Their net degeneracies 𝑁𝑠,𝛽,𝛾 are characterized, firstly, by the 𝑆𝑂(2) spin 𝑠 whose origin is most clearly seen from the M-theory perspective [5, 41]. Secondly, they depend on two-cycles 𝛽 wrapped by open M2-branes, as well as one-cycles 𝛾 on which these M2-branes can end (In case of 𝑁 D4-branes wrapping the same Lagrangian cycle, these states would additionally arise in representations 𝑅 of 𝑈(𝑁) [5]. In case of a single brane, this reduces to 𝑈(1), and such a dependence can be reabsorbed into a parameter specifying a choice of 𝛾.).

Lifting this system to M-theory, we obtain a background of the form Taub-NUT×Calabi-Yau×𝑆1, with the additional D4-brane promoted to M5-brane. This M5-brane wraps the Lagrangian submanifold 𝐿 inside Calabi-Yau, the time circle 𝑆1, and +×𝑆1TN inside the Taub-NUT space. A part of this Lagrangian 𝐿 is a torus 𝑇2=𝑆1TN×𝑆1, which will lead to some modular properties of the BPS counting functions: this modularity will be manifest in one chamber, where the open topological string amplitude will be completed to the product of 𝜃 functions. This M5-brane also breaks the 𝑆𝑂(4) spatial symmetry down to 𝑆𝑂(2)×𝑆𝑂(2). We denote the spins associated to both 𝑆𝑂(2) factors, respectively, by 𝜎 and 𝜎, and the degeneracies of particles with such spins by 𝑁𝜎,𝜎𝛽,𝛾. In addition to closed Kähler parameters 𝑄=𝑒𝑇, let us also introduce open ones related to discs wrapped by M2-branes 𝑧=𝑒𝑑. The real and imaginary parts of 𝑇 encode, respectively, the sizes of two-cycles 𝛽 and the value of the 𝐵-field through them. The real and imaginary parts of 𝑑 encode, respectively, sizes of the discs and holonomies of the gauge fields around them. Similarly as in the closed string case, to get nontrivial ensemble of mutually supersymmetric states, we set the real parts of 𝑇 and 𝑑 to zero, and consider nontrivial imaginary parts.

From the M-theory perspective, we are interested in counting the net degeneracies of M2-branes ending on this M5-brane 𝑁𝜎,𝛽,𝛾=𝜎(1)𝜎𝑁𝜎,𝜎𝛽,𝛾.(2.9) In the remaining three-dimensional space, in the 𝑅 limit, the M2-branes ending on the M5-brane are represented by a gas of free particles. These particles have excitations in 2 which we identify with the 𝑧1-plane. To each such BPS particle, similarly as in the closed string case discussed above and in [23, 40], we can associate a holomorphic field Φ𝑧1=𝑙𝛼𝑙𝑧𝑙1.(2.10) The modes of this field create states with the intrinsic spin 𝑠 and the orbital momentum 𝑙 in the 2 plane. The derivation of the BPS degeneracies relies on the identification of this total momentum 𝜎+𝑙 in the 𝑅 limit, with the Kaluza-Klein modes associated to the rotations along 𝑆1TN for the finite 𝑅, following the five-dimensional discussion in [40, 42].

The BPS generating functions we are after are given by a trace over the Fock space built by the oscillators of the second quantized field Φ(𝑧1) and restricted to the states which are mutually supersymmetric. In such a trace, each oscillator from (2.10) gives rise to one factor of the form (1𝑞𝑠𝜎+𝑙1/2𝑄𝛽𝑧𝛾)±1, where the exponent ±1 corresponds to the bosonic or fermionic character of the top component of the BPS state, 𝒵openBPS=𝜎,𝛽,𝛾𝑙=11𝑞𝑠𝜎+𝑙1/2𝑄𝛽𝑧𝛾𝑁𝜎,𝛽,𝛾|||chamber,(2.11) where the product is over either both positive or both negative (𝛽,𝛾). The parameters𝑞,𝑄 and 𝑧 specify the chamber structure: the restriction to a given chamber is implemented by imposing the condition on a central charge, analogous to (2.5), 𝑞𝑠𝜎+𝑙1/2𝑄𝛽𝑧𝛾<1.(2.12) This condition in fact specifies a choice of both closed and open chambers. The walls of marginal stability between chambers correspond to subspaces where, for some oscillator, the above product becomes 1, and then the contribution from such an oscillator drops out from the BPS generating function.

Similarly as in the closed string case, the above degeneracies can be related to open topological string amplitudes, rewritten in [5] in the form 𝒵opentop=exp𝑛=1𝜎𝛽,𝛾>0𝑁𝜎,𝛽,𝛾𝑞𝑠𝑛𝜎𝑄𝑛𝛽𝑧𝑛𝛾𝑛𝑞𝑠𝑛/2𝑞𝑠𝑛/2,(2.13) with integer Ooguri-Vafa invariants 𝑁𝜎,𝛽,𝛾 (In case of 𝑁 D4-branes wrapping a Lagrangian cycle, this structure is again more complicated, because the states in 3 arise in representations of 𝑈(𝑁) [5]. This requires replacing the factor 𝑧𝑛𝛾 by the sum 𝑅Tr𝑅𝑉𝑛 of traces in all possible representations 𝑅 of this 𝑈(𝑁) of the matrix 𝑉 encoding holonomies of the gauge fields. For simplicity we restrict here to the simplest case.). This formula represents in fact a series of quantum dilogarithms 𝐿𝑧,𝑞𝑠=exp𝑛>0𝑧𝑛𝑛𝑞𝑠𝑛/2𝑞𝑠𝑛/2=𝑛=11𝑧𝑞𝑠𝑛1/2(2.14) and can be written in the product form 𝒵opentop(𝑄,𝑧)=𝜎𝛽,𝛾>0𝑛=11𝑄𝛽𝑧𝛾𝑞𝑠𝜎+𝑛1/2𝑁𝜎,𝛽,𝛾.(2.15) Comparing with (2.11) we conclude that the BPS counting functions take form of the modulus square of the open topological string amplitude 𝒵openBPS=𝒵opentop(𝑄,𝑧)𝒵opentop𝑄1,𝑧1||chamber.(2.16)

Similarly as in the closed string case, there are also a few particularly interesting chambers to consider. For example, in the extreme chamber corresponding to Im 𝑇, Im 𝑑0, the trace is performed over the full Fock space and yields the modulus square of the open topological string partition function. In this case, the quantum dilogarithms arise in pairs, which (using the Jacobi triple product identity) combine to the modular function 𝜃3/𝜂; in consequence, the total BPS generating function is modular and expressed as a product of such functions.

2.2. Crystal Interpretation

Closed BPS generating functions (2.4) turn out to be generating functions of statistical models of crystals, when parameters relevant for both interpretations are appropriately matched. Physical reasons for such relations have been given in [8, 21, 22], and mathematical interpretation arose from works [9, 14, 15]. Such crystal interpretation arises also from the fermionic formulation [16, 17], as we will review below. These crystals, in a more intricate way [27], encode also open BPS generating functions (2.11). However, before discussing details of all these constructions, in this introductory section we present crystal models for two simplest toric Calabi-Yau manifolds, that is, 3 and conifold.

3 is the simplest Calabi-Yau manifold. It has no compact two-cycles, so relevant BPS states are bound states of arbitrary number of D0-branes with a single D6-brane wrapping entire 3. Their generating function is therefore expressed in terms of a single parameter 𝑞𝑠=𝑒𝑔𝑠. There is just a single nonzero Gopakumar-Vafa invariant 𝑁0𝛽=0=1, and as follows from (2.4) this generating function coincides with the so-called MacMahon function 𝒵BPS=𝑙=111𝑞𝑙𝑠𝑙𝑞=𝑀𝑠.(2.17)

On the other hand, the MacMahon function is a generating function of plane partitions, that is, three-dimensional generalization of Young diagrams. These plane partitions represent the simplest three-dimensional crystal model, namely, they can be identified with stacks of unit cubes filling the positive octant of 3 space, as shown in Figure 1. A unit cube located in position (𝐼,𝐽,𝐾) can evaporate from this crystal only if all other cubes with coordinates (𝑖𝐼,𝑗𝐽,𝑘𝐾) are already missing. A plane partition 𝜋 is weighted by the number of boxes it consists of |𝜋|, with a weight 𝑞 associated to a single box, so indeed 𝑍=𝜋𝑞|𝜋|=𝑙=0𝑝(𝑙)𝑞𝑙=1+𝑞+3𝑞2+6𝑞3+13𝑞4+=𝑀(𝑞),(2.18) where 𝑝(𝑙) is the number of plane partitions which consist of 𝑙 cubes. Therefore, plane partition generating function coincides with the BPS counting function 𝑍=𝒵BPS when a simple identification 𝑞𝑠=𝑞(2.19) is made. From (2.6), it follows that the topological string partition function for 3 is given by the square root of the MacMahon function 𝒵top𝑞=𝑀𝑠1/2,(2.20) which is indeed true. The relevance of the MacMahon function for 3 geometry was noticed for the first time in [3], and a statistical model interpretation of this result was proposed in [7].

357016.fig.001
Figure 1: Plane partitions represent melting crystal configurations of 3.

The conifold provides another simple, yet nontrivial example of toric Calabi-Yau manifold. It consists of two 3 patches glued int𝒪(1)𝒪(1)1, and it has one Kähler class representing 1, parametrized by 𝑄=𝑒𝑇. This class can be wrapped by D2-branes, which bind with D0-branes to an underlying D6-brane and give rise to BPS states in low energy theory. In this case, there is already a nontrivial structure of chambers and walls, which was analyzed in [14, 18, 20, 21]. This structure is consistent with M-theory derivation discussed in Section 2.1. The generating functions of D6-D2-D0 bound states are parametrized by 𝑄 and 𝑞𝑠, and therefore corresponding crystal models consist of two-colored three-dimensional partitions. The Kähler moduli space consists of several infinite countable sets of chambers, and in each chamber relevant crystal configurations take form of so-called pyramid partitions. These partitions are infinite or finite (resp. for positive and negative 𝑅 in (2.1)) and their size depends on the value of the 𝐵-field. This size changes discretely and the pyramid is enlarged when the value of the 𝐵-field crosses integer numbers, which changes the chamber in the moduli space, as explained in Section 2.1. Examples of such infinite pyramid partitions are given in Figure 2, and finite ones in Figure 3.

fig2
Figure 2: Infinite pyramids with one and four balls in the top row, with generating functions given, respectively, by 𝑍pyramid0 and 𝑍pyramid3.
fig3
Figure 3: Finite pyramids with 𝑚=1,2,3 stones in the top row (resp., (a), (b), and (c)), whose generating functions are given by 𝑍pyramid𝑚+1 (note that 𝑍pyramid1=1 corresponds to an empty pyramid corresponding to the pure D6-brane).

To write down explicitly BPS generating functions for the conifold in various chambers, we can take advantage of their relation to the topological string amplitude (2.6). The topological string partition function in this case reads 𝒵conifoldtop𝑞(𝑄)=𝑀𝑠𝑘11𝑄𝑞𝑘𝑠𝑘,(2.21) with the MacMahon function defined in (2.17). From this topological string partition function we can read off Gopakumar-Vafa invariants [3, 4] 𝑁0𝛽=0=2,𝑁0𝛽=±1=1.(2.22)

Using the relation (2.6), we can now present conifold closed BPS generating functions in several sets of chambers. In the first set of chambers, we consider 𝑅>0 and positive 𝐵]𝑛,𝑛+1[ (for 𝑛0). Firstly, for small 𝐵, there is so-called noncommutative chamber discussed first by Szendrői [14], which corresponds to 𝑛=0. In this case, the pyramid crystal has just a single ball in the top row, as in Figure 2(a), and the BPS generating function is given by the square of the topological amplitude. On the other hand, for large 𝐵, that is 𝑛, we reach commutative chamber in which the length of the top row extends to infinity. In this case, the BPS generating function agrees, up to a single factor of MacMahon function, with the topological string amplitude. In between, there are chambers with 𝑛+1 balls in the top row, for which 𝒵conifold𝑛𝑞=𝑀𝑠2𝑘11𝑄𝑞𝑘𝑠𝑘𝑘𝑛+11𝑄1𝑞𝑘𝑠𝑘.(2.23) These BPS generating functions are related to pyramid generating functions with two colors 𝑞0 and 𝑞1 upon the identification (which generalizes (2.19) in 3 case) 𝒵conifold𝑛chambers𝑞𝑠=𝑞0𝑞1,𝑄=𝑞𝑛𝑠𝑞1.(2.24) Indeed, with this identification, the above counting functions agree with those of two-colored pyramid crystals with 𝑛+1 yellow balls in its top row 𝑍pyramid𝑛𝑞0,𝑞1𝑞=𝑀0𝑞12𝑘𝑛+11+𝑞𝑘0𝑞1𝑘+1𝑘𝑛𝑘11+𝑞𝑘0𝑞1𝑘1𝑘+𝑛.(2.25)

In the second set of chambers, we have 𝑅<0 and positive 𝐵]𝑛1,𝑛[ (for 𝑛1). It extends between the core region with a single D6-brane (2.8) and the chamber characterized by so-called Pandharipande-Thomas invariants (for the flopped geometry, or equivalently for anti-M2-branes). The BPS generating functions read 𝒵conifold𝑛=𝑛1𝑗=1𝑞1𝑗𝑠𝑄𝑗.(2.26) The corresponding statistical models were shown in [16, 18, 21] to correspond to finite pyramids with 𝑛1 stones in the top row, as shown in Figure 3. In this case, the generating functions of such partitions are equal to 𝑍pyramid𝑛𝑞0,𝑞1=𝑛1𝑗=11+𝑞0𝑛𝑗𝑞1𝑛𝑗1𝑗.(2.27) The equality 𝒵conifold𝑛𝑍pyramid𝑛 arises upon an identification 𝒵conifold𝑛chambers𝑞𝑠1=𝑞0𝑞1,𝑄=𝑞𝑛𝑠𝑞1.(2.28) There are two other sets of chambers characterized by the negative value of the 𝐵-field, for which BPS generating functions are completely analogous to those given above.

Above, we presented just the simplest examples of crystal models. Using fermionic formulation presented below, one can find other crystal models for arbitrary toric geometry without compact four cycles. Let us also mention that those models can be equivalently expressed in terms dimers. In particular, the operation of enlarging the crystal, as in the conifold pyramids, corresponds to so-called dimer shuffling [15]. Dimers are also closely related to a formulation using quivers and associated potentials, which underlies physical derivations in [21, 22].

3. A Little Background—Free Fermions and Matrix Models

In this section, we introduce some mathematical background on which the main results presented in this paper rely. In Section 3.1, we start with a brief presentation of toric Calabi-Yau manifolds and introduce the notation which we use in what follows. In Section 3.2, we introduce free fermion formalism. In Section 3.3, we introduce basics of matrix model formalism. Our presentation is necessarily brief, and for more detailed introduction we recommend many excellent reviews on each of those topics.

3.1. Toric Calabi-Yau Threefolds

Some introductory material on toric Calabi-Yau manifolds, from the perspective relevant for mirror symmetry and topological string thoery, can be found, for example, in [39]. In this section, our presentation is brief and mainly sets up the notation. Toric Calabi-Yau threefolds arise as the quotient of 𝜅+3, possibly with a discrete set of points deleted, by the action of ()𝜅 with certain weights. The simplest toric threefold is 3, which corresponds to the trivial choice 𝜅=0. The resolved conifold, which we already discussed in Section 2.2, corresponds to 𝜅=1 and a choice of weights (1,1,1,1), which represent a local bundle 𝒪(1)𝒪(1)1. The structure of each toric three-fold can be encoded in a two-dimensional diagram built from trivalent vertices. Finite intervals joining two adjacent vertices represent local 1 neighborhood inside the manifold. Equivalently, one can consider dual graphs. Examples of toric diagrams and their duals for 3, conifold and resolution of 3/2 singularity are given in Figure 4 (the notation Γ± at each vertex will be explained in what follows).

357016.fig.004
Figure 4: Toric graphs for 3, conifold and resolution of 3/2.

A closed loop in a toric diagram represents a compact four cycles in the geometry. As follows from the reasoning in Section 2.1, in the context of BPS counting, we are forced to restrict considerations to manifolds which do not have such four cycles. Apart from a few special cases, there is an infinite class of such geometries whose dual diagrams arise from a triangulation, into triangles of area 1/2, of a long rectangle or a strip of height 1. A toric diagram arises as a dual graph to such a triangulation. From each vertex in such a toric diagram, one vertical line extends to infinity and crosses either the upper or the lower edge of the strip. Two such consecutive lines can emanate either in the same or in the opposite direction, respectively, when they are the endpoints of an interval representing 1 with local 𝒪(2)𝒪 or 𝒪(1)𝒪(1) neighborhood. An example of a generic diagram of this kind is shown in Figure 8.

Let us denote independent 1's, starting from the left end of the strip, from 1 to 𝑁, and introduce corresponding Kähler parameters 𝑄𝑖=𝑒𝑇𝑖, 𝑖=1,,𝑁. Moreover, to each toric vertex we associate a type 𝑡𝑖=±1, so that 𝑡𝑖+1=𝑡𝑖 if the local neighborhood of 1 (represented by an interval between vertices 𝑖 and 𝑖+1) is 𝒪(2)𝒪; if this neighborhood is of 𝒪(1)𝒪(1) type, then 𝑡𝑖+1=𝑡𝑖. The type of the first vertex we fix as 𝑡1=+1. In Figures 4 and 8, these types are denoted by and . The types 𝑡𝑖 will be used much in the construction of fermionic states in Section 4.2.

As explained in Section 2.1, the BPS generating functions can be expressed in terms (the instanton part) of topological string amplitudes. For the above class of geometries, arising from a triangulation of a strip, these amplitudes read 𝒵top𝑄𝑖𝑞=𝑀𝑠(𝑁+1)/2𝑙=11𝑖<𝑗𝑁+11𝑞𝑙𝑠𝑄𝑖𝑄𝑖+1𝑄𝑗1(𝑡𝑖𝑡𝑗)𝑙.(3.1)

3.2. Free Fermion Formalism

Formalism of free fermions in two dimensions is well known [43, 44] and ubiquitous in literature on topological strings and crystal melting [7, 15, 15, 37, 45]. The main purpose of this section is therefore to set up the notation which we will follow in the remaining parts of this paper.

The states in the free fermion Fock space are created by the (anticommuting) modes of the fermion field 𝜓(𝑧)=𝑘𝜓𝑘+1/2𝑧𝑘1,𝜓(𝑧)=𝑘𝜓𝑘+1/2𝑧𝑘1,𝜓𝑘+(1/2),𝜓𝑙1/2=𝛿𝑘,𝑙(3.2) on the vacuum state |0. There is one-to-one map between such fermionic states ||𝜇=𝑑𝑖=1𝜓𝑎𝑖1/2𝜓𝑏𝑖1/2||0,with𝑎𝑖=𝜇𝑖𝑖,𝑏𝑖=𝜇𝑡𝑖𝑖,(3.3) and two-dimensional partitions 𝜇=(𝜇1,𝜇2,,𝜇𝑙), as shown in Figure 5. The modes 𝛼𝑚 of the bosonized field 𝜕𝜙=𝜓(𝑧)𝜓(𝑧) satisfy the Heisenberg algebra [𝛼𝑚,𝛼𝑛]=𝑛𝛿𝑚,𝑛.

357016.fig.005
Figure 5: Relation between Young diagrams and states in the Fermi sea.

We introduce vertex operators Γ±(𝑥)=𝑒𝑛>0(𝑥𝑛/𝑛)𝛼±𝑛,Γ±(𝑥)=𝑒𝑛>0((1)𝑛1𝑥𝑛/𝑛)𝛼±𝑛,(3.4) which act on fermionic states |𝜇 corresponding to partitions 𝜇 as [15, 43, 44] Γ(||𝑥)𝜇=𝜆𝜇𝑥|𝜆||𝜇|||𝜆,Γ+(||𝑥)𝜇=𝜆𝜇𝑥|𝜇||𝜆|||Γ𝜆,(3.5)||(𝑥)𝜇=𝜆𝑡𝜇𝑡𝑥|𝜆||𝜇|||𝜆,Γ+||(𝑥)𝜇=𝜆𝑡𝜇𝑡𝑥|𝜇||𝜆|||𝜆.(3.6) The interlacing relation between partitions is defined as 𝜆𝜇𝜆1𝜇1𝜆2𝜇2𝜆3.(3.7)

The operator Γ is the inverse of Γ with negative argument. These operators satisfy commutation relations Γ+(𝑥)Γ1(𝑦)=Γ1𝑥𝑦(𝑦)Γ+Γ(𝑥),(3.8)+(𝑥)Γ1(𝑦)=Γ1𝑥𝑦(𝑦)Γ+Γ(𝑥),(3.9)+(𝑥)Γ(𝑦)=(1+𝑥𝑦)Γ(𝑦)Γ+Γ(𝑥),(3.10)+(𝑥)Γ(𝑦)=(1+𝑥𝑦)Γ(𝑦)Γ+(𝑥).(3.11)

We also introduce various colors 𝑞𝑔 and the corresponding operators 𝑄𝑔 (a hat is to distinguish them from Kähler parameters 𝑄𝑖) 𝑄𝑔||𝜆=𝑞𝑔|𝜆|||𝜆.(3.12) These operators commute with vertex operators up to rescaling of their arguments Γ+𝑄(𝑥)𝑔=𝑄𝑔Γ+𝑥𝑞𝑔,Γ+𝑄(𝑥)𝑔=𝑄𝑔Γ+𝑥𝑞𝑔,𝑄(3.13)𝑔Γ(𝑥)=Γ𝑥𝑞𝑔𝑄𝑔,𝑄𝑔Γ(𝑥)=Γ𝑥𝑞𝑔𝑄𝑔.(3.14)

3.3. Matrix Models

In matrix model theory, or theory of random matrices, one is interested in properties of various ensembles of matrices. Excellent reviews of random matrix theory can be found for example in [46] or, in particular in the context of topological string theory, in [47]. In matrix model theory, one typically considers partition functions of the form 𝑍=𝒟𝑈𝛼𝑒(1/𝑔𝑠)Tr𝑉(𝑈),(3.15) where 𝑉=𝑉(𝑈) is a matrix potential and 𝒟𝑈 is a measure over a set of matrices of interest 𝑈 of size 𝑁. Typically it is not possible to perform the above integral; however, special techniques allow to determine its formal 1/𝑁 expansion. These techniques culminated with the formalism of the topological expansion of Eynard and Orantin [48] which, in principle, allows to determine entire 1/𝑁 expansion of the partition function recursively. This solution is determined by the behavior of matrix eigenvalues, whose distribution among the minima of the potential, in the continuum limit, determines one-dimensional complex curve, so-called spectral curve. The spectral curve is also encoded in the leading 1/𝑁 expansion of the so-called resolvent, which is defined as the expectation value 𝜔(𝑥)=Tr(1/(𝑥𝑈)) computed with respect to the measure (3.15).

In the context of BPS counting and topological strings, unitary ensembles of matrices of infinite size arise. In this case, the matrix model simplifies to the integral over eigenvalues 𝑢𝛼, with a measure which takes form of the unitary Vandermonde determinant 𝒟𝑈=𝛼𝑑𝑢𝛼𝛼<𝛽||𝑧𝛼𝑧𝛽||2,𝑧𝛼=𝑒𝑖𝑢𝛼.(3.16) The issue of infinite matrices is a little subtle; however, it can be taken care of by considering matrices of large but finite size 𝑁, and subsequently taking 𝑁 limit. For finite 𝑁, one can find the resolvent, and in consequence the spectral curve, using a standard technique of so-called Migdal integral. This requires redefining 𝑉 to the standard Vandermonde form [29, 47], as well as introducing 't Hooft coupling 𝑇𝑉𝑉+𝑇log𝑧,𝑇=𝑁𝑔𝑠.(3.17) The form of the Migdal integral depends on the number of cuts into which eigenvalues condense in large 𝑁 limit, and this number of cuts determines the genus of the spectral curve. In our context, only single-cut situations will arise, for which the spectral curve has genus zero. In this case, the Migdal integral determines the resolvent as 1𝜔(𝑝)=2𝑇𝑑𝑧𝜕2𝜋𝑖𝑧𝑉(𝑧)𝑝𝑧(𝑝𝑎)(𝑝𝑏)(𝑧𝑎)(𝑧𝑏),(3.18) so that the integration contour encircles counter-clockwise the endpoints of the cut 𝑎 and 𝑏. A proper asymptotic behavior of the resolvent is imposed by the condition lim𝑝1𝜔(𝑝)=𝑝.(3.19) Then the spectral curve is determined as a surface on which the resolvent is unambiguously defined, that is, it is given by an (exponential) rational equation automatically satisfied by 𝑝 and 𝜔(𝑝). There is also an important consistency condition for the resolvent: when computed on the opposite sides of the cut 𝜔(𝑝)±, it is related to the potential as 𝜔+(𝑝)+𝜔𝜕(𝑝)=𝑝𝑉(𝑝)𝑇.(3.20) On the other hand, a difference of these values of the resolvent on both sides of the cut provides eigenvalue density 𝜌(𝑝)=𝜔+(𝑝)𝜔(𝑝).(3.21)

It has been observed in several contexts that topological strings on toric manifolds can be related to matrix models, whose spectral curves take form of the so-called mirror curves. Mirror curves arise for manifolds which are mirror to toric Calabi-Yau manifolds [39, 45]. For toric manifolds, their mirror manifolds are determined by the following equation embedded in four-dimensional complex space: 𝑧1𝑧2=𝐻(𝑥,𝑦).(3.22) The mirror curve is the zero locus of 𝐻(𝑥,𝑦), that is, it is given as 𝐻(𝑥,𝑦)=0. More precisely, 𝑥,𝑦 are variables, and it is often convenient to represent them in the exponential form 𝑥=𝑢𝑢, 𝑦=𝑒𝑣, with 𝑢,𝑣. For example, for 3 and the conifold they take the following form: 𝐻3(𝑥,𝑦)=𝑥+𝑦+𝑥𝑦=0,𝐻conifold(𝑥,𝑦)=𝑥+𝑦+𝑥𝑦+𝑄𝑥2=0,(3.23) where 𝑄 encodes the Kähler parameter of the conifold. Schematically mirror curves arise from thickening edges of the toric graphs, as shown in Figure 6.

357016.fig.006
Figure 6: Toric diagrams for 3 and conifold and corresponding mirror curves.

One of the first relations between topological strings for toric manifolds and matrix models was encountered in [49, 50], where it was shown that the spectral curve of a unitary matrix model with a Gaussian (i.e., quadratic) potential agrees with the above mirror curve 𝐻conifold(𝑥,𝑦)=0 in (3.23), with ’t Hooft coupling 𝑇=𝑔𝑠𝑁 encoded in 𝑄=𝑒𝑇. At the same time, it was shown that the matrix model partition function reproduces the topological string partition function. More recently these ideas became important in view of the remodeling conjecture [51], which states that the solution to loop equations in the form found by Eynard and Orantin [48], applied to the mirror curve, reproduces topological string partition functions. The method of [48] works for arbitrary curves, not necessarily originating from matrix models. Nonetheless, it is indeed possible to construct matrix models whose partition functions do reproduce topological string amplitudes, and whose spectral curves coincide with appropriate mirror curves [29, 30, 5256].

One of our aims is to provide matrix model interpretation of BPS counting. It is natural to expect such an interpretation in view of an intimate relation between BPS counting and topological string theory discussed in Section 2.1, and the above-mentioned relations between topological strings and matrix models. As we will see in what follows, there are indeed unitary matrix models which naturally arise in the context of BPS counting and its fermionic formulation. Among the others, our task will be to analyze them using the above-mentioned Migdal method.

4. Fermionic Formulation of BPS Counting Functions

Having introduced all the ingredients above, we are now ready to present fermionic formulation of BPS counting. To start with, in Section 4.1 we present the idea of such a formulation in the simplest example of 3. In Section 4.2, we introduce a general fermionic formalism, and in Section 4.3 we provide its crystal interpretation. We illustrate the use of our formalism in Section 4.4 revisiting 3 example, as well as in explicit case of 3/𝑁, and conifold geometry.

4.1. The Idea and 3 Example

As explained in Section 2.2, the generating function of bound states of D0-branes to a single D6-brane is given by the MacMahon function, and the corresponding crystal model takes form of the counting of plane partitions [7]. Let us slice each such plane partition by a set of parallel planes, as shown in Figure 7. In this way on each slice, we obtain a two-dimensional partition 𝜇, and it is not hard to see that each two neighboring partitions satisfy the interlacing condition (3.7). Recalling that such a condition arises if we apply Γ±(1) operators (3.5) to partition states, we conclude that a set of all plane partitions can be built, slice by slice, by acting with infinite sequence of Γ±(1) on the vacuum. To count each slice 𝜇 with appropriate weight 𝑞|𝜇| we also need to apply weight operator 𝑄 defined in (3.12). Therefore, the generating function of plane partitions can be represented as followsΩ𝑍=+Ω||0𝑄Γ+(1)𝑄Γ+(1)𝑄Γ+(1)𝑄Γ(1)𝑄Γ(1)𝑄Γ||||(1)𝑄0=0Γ+𝑞2Γ+(𝑞)Γ+(1)Γ(𝑞)Γ𝑞2Γ𝑞3||=0𝑙1,𝑙2=111𝑞𝑙1+𝑙21=𝑀(𝑞).(4.1) In the first line, we implicitly introduced two states Ω+| and |Ω, defined by an infinite sequence of Γ+ (resp. Γ) operators, interlaced with weight operators 𝑄 and acting on the vacuum. To confirm that this correlator indeed reproduces the MacMahon function, the second line can be reduced to the final infinite product using commutation relations (3.8) and (3.14). We can also represent insertions of Γ±(1) operators graphically by arrows, so that the above computation can be represented as in Figure 7(b).

357016.fig.007
Figure 7: Slicing of a plane partition (a) into a sequence of interlacing two-dimensional partitions (b). A sequence of Γ± operators in (4.1) which create two-dimensional partitions is represented by arrows inserted along two axes. Directions of arrows represent interlacing condition on partitions. We reconsider this example from a new viewpoint in Figure 10.
357016.fig.008
Figure 8: Toric Calabi-Yau manifolds represented by a triangulation of a strip. There are 𝑁 independent 1's with Kähler parameters 𝑄𝑖=𝑒𝑇𝑖, and 𝑁+1 vertices to which we associate Γ and Γ operators represented respectively by and signs. Yellow intervals, which connect vertices with opposite signs, represent 𝒪(1)𝒪(1)1 local neighborhoods. Red intervals, which connect vertices with the same signs, represent 𝒪(2)𝒪1 local neighborhoods. The first vertex on the left is chosen to be .

In what follows, we present a formalism which allows to generalize this computation to a large class of chambers, for arbitrary toric geometry without compact four cycles.

4.2. Toric Geometry and Quantization

We wish to reformulate BPS counting in the fermionic language in a way in which we associate to each toric manifold a fermionic state, such that the BPS generating function can be expressed as an overlap of two such states, generalizing 3 case (4.1). At the same time, the construction of such a fermionic state is supposed to encode the structure of the underlying crystal model (generalizing plane partitions in Figure 7). An important difference between 3 and other geometries is the existence of many Kähler moduli and correspondingly many chambers, for which BPS generating functions change according to wall-crossing formulas. To take care of these changes in the fermionic formalism, we need to introduce special wall-crossing operators.

4.2.1. Toric Geometry and Fermionic Operators

In what follows we use the notation introduced in Section 3.1; in particular to each vertex of the toric diagram we associate its type 𝑡𝑖=±1, see also Figure 8. We start with a construction of fermionic states associated to a given toric Calabi-Yau manifold (without compact four-cycles). First we need to introduce several operators which are building blocks of such states. The structure of these operators is encoded in the toric diagram of a given manifold. Namely, these operators are given by a string of 𝑁+1 vertex operators Γ𝑡𝑖±(𝑥) (defined in (3.4)) which are associated to the vertices of the toric diagram; the type 𝑡𝑖 determines the type of a vertex operator as Γ𝑡𝑖±=+1(𝑥)=Γ±(𝑥),Γ𝑡𝑖±=1(𝑥)=Γ±(𝑥).(4.2) In addition the string of operators Γ𝑡𝑖±(𝑥) is interlaced with 𝑁+1 operators 𝑄𝑖 representing colors 𝑞𝑖, for 𝑖=0,1,,𝑁. Operators 𝑄1𝑄,,𝑁 are associated to 1 in the toric diagram, and there is an additional 𝑄0. We also define 𝑄𝑄=0𝑄1𝑄𝑁,𝑞=𝑞0𝑞1𝑞𝑁.(4.3) Therefore, the upper indices of Γ𝑡𝑖±(𝑥) and a choice of colors of the operators which we introduce below are specified by the data of a given toric manifold. As we will see, a sequence of lower indices ± is determined by the chamber we are going to consider.

Now we can associate several operators to a given toric manifold. Firstly, we define 𝐴±(𝑥)=Γ𝑡1±𝑄(𝑥)1Γ𝑡2±𝑄(𝑥)2Γ𝑡𝑁±𝑄(𝑥)𝑁Γ𝑡𝑁+1±𝑄(𝑥)0.(4.4) Commuting all 𝑄𝑖's using (3.14), we also define the following operators: 𝐴+𝑄(𝑥)=1𝐴+(𝑥)=Γ𝑡1+(𝑥𝑞)Γ𝑡2+𝑥𝑞𝑞1Γ𝑡3+𝑥𝑞𝑞1𝑞2Γ𝑡𝑁+1+𝑥𝑞𝑞1𝑞2𝑞𝑁,𝐴(𝑥)=𝐴(𝑄𝑥)1=Γ𝑡1(𝑥)Γ𝑡2𝑥𝑞1Γ𝑡3𝑥𝑞1𝑞2Γ𝑡𝑁+1𝑥𝑞1𝑞2𝑞𝑁.(4.5)

In addition, we define the above-mentioned wall-crossing operators 𝑊𝑝Γ(𝑥)=𝑡1𝑄(𝑥)1Γ𝑡2𝑄(𝑥)2Γ𝑡𝑝𝑄(𝑥)𝑝Γ𝑡𝑝+1+𝑄(𝑥)𝑝+1Γ𝑡𝑁+𝑄(𝑥)𝑁Γ𝑡𝑁+1+𝑄(𝑥)0,𝑊𝑝Γ(𝑥)=𝑡1+𝑄(𝑥)1Γ𝑡2+𝑄(𝑥)2Γ𝑡𝑝+𝑄(𝑥)𝑝Γ𝑡𝑝+1𝑄(𝑥)𝑝+1Γ𝑡𝑁𝑄(𝑥)𝑁Γ𝑡𝑁+1𝑄(𝑥)0.(4.6) Here the order of Γ and Γ is the same as for 𝐴± operators, and the difference is that now there are subscripts on first 𝑝 operators and ± on the remaining ones.

We often use a simplified notation when the argument of the above operators is 𝑥=1𝐴±𝐴±(1),𝐴±𝐴±(1),𝑊𝑝𝑊𝑝(1),𝑊𝑝𝑊𝑝(1).(4.7)

4.2.2. Fermionic Formulation and Quantization

Above we associated operators 𝐴± to each toric geometry with a strip-like toric diagram. From these operators, we can build the following states in the Hilbert space of a free fermion : ||Ω±,(4.8) which we define as follows: Ω+||||=0𝐴+(1)𝐴+(1)𝐴+||(1)=0𝐴+𝑞2𝐴+(𝑞)𝐴+||Ω(1),=𝐴(1)𝐴(1)𝐴(||1)0=𝐴(1)𝐴(𝑞)𝐴𝑞2||0.(4.9) These states encode the full instanton part of the topological string amplitudes. Namely, as shown in [16], Ω𝑍=+Ω(4.10) is equal to the BPS partition function 𝒵 in the noncommutative chamber ||𝒵𝑍=𝒵top||2𝒵top𝑄𝑖𝒵top𝑄𝑖1,(4.11) where 𝒵top(𝑄𝑖) is given in (3.1). The above equality holds under the following identification between 𝑞𝑖 parameters (which enter the definition of |Ω±) and physical parameters 𝑄𝑖=𝑒𝑇𝑖 and 𝑞𝑠=𝑒𝑔𝑠: 𝑞𝑖=𝑡𝑖𝑡𝑖+1𝑄𝑖,𝑞𝑠=𝑞𝑞0𝑞1𝑞𝑁.(4.12) We will provide a proof of (4.10) in Section 6.1.1 in a more general setting of refined invariants.

The states |Ω± have nontrivial structure and encode the information about the noncommutative chamber. It turns out that the fermionic vacuum |0 itself also encodes some interesting information. We recall that there is another extreme chamber representing just a single BPS state represented by the D6-brane with no other branes bound to it. This multiplicity 1 can be understood as 𝒵=𝑍=00=1,(4.13) and as we will see below, starting from this expression we can use wall-crossing operators to construct BPS generating functions in an infinite family of other chambers.

4.2.3. Other Chambers and Wall-Crossing Operators

In the previous, section we associated to toric manifolds the states |Ω±, whose overlap reproduces the BPS generating function in the noncommutative chamber (4.10). Now we wish to extend this formalism to other chambers. As discussed in Section 2.1, in a given chamber, the allowed bound states we wish to count must have positive central charge (2.1) 1𝑍(𝑅,𝐵)=𝑅(𝑛+𝛽𝐵)>0.(4.14) Firstly, the information about 𝑅 and 𝐵 must be encoded in the fermionic states which we wish to construct. It turns out that the choice of positive or negative 𝑅 is encoded in the choice of the ground state ||Ω𝑅>0±||,𝑅<00,(4.15) which generalizes the extreme cases (4.10) and (4.13).

On the other hand, the value of the field 𝐵 is encoded in the insertion of additional wall-crossing operators, such as those defined in (4.6). In particular, these two types of operators are sufficient if we wish to consider only these chambers, which correspond to a flux of the 𝐵-field through only one, but arbitrary 1 in the manifold. For simplicity below, we consider only this set of chambers. Denoting this 1 as 𝑝, it can be shown that insertion of 𝑛 copies of operators 𝑊𝑝 or 𝑊𝑝 creates, respectively, 𝑛 positive or negative quanta of the flux through 𝑝'th 1.

Therefore, schematically, the generating functions in chambers with 𝑅>0 read 𝑍𝑛=Ω+||𝑊𝑛||Ω,(4.16) and those with 𝑅<0 read 𝑍𝑛||=0𝑊𝑛||0,(4.17) with appropriate form of wall-crossing operators. More precisely, depending on the signs of 𝑅 and 𝐵, we need to consider four possible situations, which we present below. The proofs of all statements below, corresponding to these four situations, can be found in [16].

(i) Chambers with 𝑅<0, 𝐵>0
Consider a chamber characterized by positive 𝑅 and positive 𝐵-field through 𝑝'th two-cycle ][𝑅<0,𝐵𝑛1,𝑛,for1𝑛.(4.18) The BPS partition function in this chamber contains only those factors which include 𝑄𝑝 and it reads 𝒵𝑛|𝑝=𝑛1𝑝𝑖=1𝑠=1𝑁+1𝑟=𝑝+1𝑞1𝑖𝑠𝑄𝑠𝑄𝑠+1𝑄𝑟1𝑡𝑟𝑡𝑠𝑖.(4.19) This can be expressed as the expectation value of 𝑛 wall-crossing operators 𝑊𝑝𝑍𝑛|𝑝||=0𝑊𝑝𝑛||𝒵0=𝑛|𝑝,(4.20) under the following identification of variables: 𝑄𝑝=𝑡𝑝𝑡𝑝+1𝑞𝑝𝑞𝑛𝑠,𝑄𝑖=𝑡𝑖𝑡𝑖+1𝑞𝑖for𝑖𝑝,𝑞𝑠=1𝑞.(4.21) A special case of this result is the trivial generating function (4.13) representing a single D6-brane.

(ii) Chambers with 𝑅>0, 𝐵>0
In the second case, we consider the positive value of 𝑅 and the positive flux through 𝑝'th 1][𝑅>0,𝐵𝑛,𝑛+1,for0𝑛.(4.22) Denote the BPS partition function in this chamber by 𝒵𝑛|𝑝. We find that the expectation value of 𝑛 wall-crossing operators 𝑊𝑝 in the background of |Ω has the form 𝑍𝑛|𝑝=Ω+||𝑊𝑝𝑛||Ω=𝑀(1,𝑞)𝑁+1𝑍(0)𝑛|𝑝𝑍(1)𝑛|𝑝𝑍(2)𝑛|𝑝,(4.23) where 𝑍(0)𝑛|𝑝 does not contain any factors (𝑞𝑠𝑞𝑟1)±1 which would include 𝑞𝑝, while 𝑍(1)𝑛|𝑝 contains all factors 𝑞𝑠𝑞𝑟1 which do include 𝑞𝑝, and 𝑍(2)𝑛|𝑝 contains all factors (𝑞𝑠𝑞𝑟1)1 which also include 𝑞𝑝: 𝑍(0)𝑛|𝑝=𝑙=1𝑝𝑠,𝑟+11,𝑁+1𝑡1𝑟𝑡𝑠𝑞𝑙𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙𝑡1𝑟𝑡𝑠𝑞𝑙𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙,𝑍(1)𝑛|𝑝=𝑙=1𝑝𝑠,𝑟+11,𝑁+1𝑡1𝑟𝑡𝑠𝑞𝑙+𝑛𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙,𝑍(2)𝑛|𝑝=𝑙=𝑛+1𝑝𝑠,𝑟+11,𝑁+1𝑡1𝑟𝑡𝑠𝑞𝑙𝑛𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙.(4.24) We see that the identification of variables 𝑄𝑝=𝑡𝑝𝑡𝑝+1𝑞𝑝𝑞𝑛𝑠,𝑄𝑖=𝑡𝑖𝑡𝑖+1𝑞𝑖for𝑖𝑝,𝑞𝑠=𝑞(4.25) reproduces the BPS partition function 𝒵𝑛|𝑝=𝑍𝑛|𝑝.(4.26) When no wall-crossing operator is inserted the change of variables reduces to (4.12) and we get the noncommutative Donaldson-Thomas partition function (4.11), 𝑍0|𝑝=𝒵.

(iii) Chambers with R<0, B<0
Now we consider negative 𝑅 and negative 𝐵-field ][𝑅<0,𝐵𝑛1,𝑛for0𝑛.(4.27) For such a chamber the BPS partition function reads 𝒵𝑛|𝑝=𝑛𝑝𝑖=1𝑠=1𝑁+1𝑟=𝑝+11𝑞𝑖𝑠𝑄𝑠𝑄𝑠+1𝑄𝑟1𝑡𝑟𝑡𝑠𝑖.(4.28) Now we find the expectation value of 𝑛 wall-crossing operators 𝑊𝑝 is equal to 𝑍𝑛|𝑝||=0𝑊𝑝𝑛||𝒵0=𝑛|𝑝,(4.29) under the change of variables 𝑄𝑝=𝑡𝑝𝑡𝑝+1𝑞𝑝𝑞𝑠𝑛,𝑄𝑖=𝑡𝑖𝑡𝑖+1𝑞𝑖for𝑖𝑝,𝑞𝑠=1𝑞.(4.30) Now an insertion of 𝑊𝑝 has an interpretation of turning on a negative quantum of 𝐵-field, and the redefinition of 𝑄𝑝 can be interpreted as effectively reducing 𝑡𝑝 by one unit of 𝑔𝑠. As already discussed, 𝑍0𝑝=00=1(4.31) represents a chamber with a single D6-brane and no other branes bound to it.

(iv) Chambers with 𝑅>0, 𝐵<0
In the last case, we consider positive 𝑅 and negative 𝐵][𝑅>0,0>𝐵𝑛,𝑛+1,for1𝑛.(4.32) We denote the BPS partition function in this chamber by 𝒵𝑛|𝑝. We find that the expectation value of 𝑛 operators 𝑊𝑝 in the background of |Ω± has the form 𝑍𝑛|𝑝=Ω+||𝑊𝑝𝑛||Ω=𝑀(1,𝑞)𝑁+1𝑍(0)𝑛|𝑝𝑍(1)𝑛|𝑝𝑍(2)𝑛|𝑝,(4.33) where 𝑍(0)𝑛|𝑝 does not contain any factors (𝑞𝑠𝑞𝑟1)±1 which would include 𝑞𝑝, 𝑍(1)𝑛|𝑝 contains all factors 𝑞𝑠𝑞𝑟1 which do include 𝑞𝑝, and 𝑍(2)𝑛|𝑝 contains all factors (𝑞𝑠𝑞𝑟1)1 which also include 𝑞𝑝: 𝑍(0)𝑛|𝑝=𝑙=1𝑝𝑠,𝑟+11,𝑁+1𝑡1𝑟𝑡𝑠𝑞𝑙𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙𝑡1𝑟𝑡𝑠𝑞𝑙𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙,𝑍(1)𝑛|𝑝=𝑙=𝑛𝑝𝑠,𝑟+11,𝑁+1𝑡1𝑟𝑡𝑠𝑞𝑙𝑛𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙,𝑍(2)𝑛|𝑝=𝑙=1𝑝𝑠,𝑟+11,𝑁+1𝑡1𝑟𝑡𝑠𝑞𝑙+𝑛𝑞𝑠𝑞𝑠+1𝑞𝑟1𝑡𝑟𝑡𝑠𝑙.(4.34) Under the change of variables, 𝑄𝑝=𝑡𝑝𝑡𝑝+1𝑞𝑝𝑞𝑠𝑛1,𝑄𝑖=𝑡𝑖𝑡𝑖+1𝑞𝑖,for𝑖𝑝,𝑞𝑠=𝑞.(4.35) This reproduces the BPS partition function 𝒵𝑛|𝑝=𝑍𝑛|𝑝.(4.36) We note that both 𝑍1|𝑝 with the above change of variables, as well as 𝑍0|𝑝 given in (4.23) with a different change of variables in (4.25), lead to the same BPS generating function 𝒵 which corresponds to the noncommutative Donaldson-Thomas invariants.

4.3. Crystal Melting Interpretation

In the previous section, we found a free fermion representation of D6-D2-D0 generating functions. The fermionic correlators which reproduce BPS generating functions automatically provide melting crystal interpretation of these functions [16], generalizing models of plane partitions (for 3) or pyramid partitions (for the conifold), presented in Section 2.2. These crystals are also equivalent to those found in [17, 22].

The crystal interpretation is a consequence of the fact that all operators used in the construction of states |Ω±, as well as the wall-crossing operators, are built just from vertex operators Γ± and Γ± with argument 1, and color operators 𝑄𝑖. As follows from (3.5) and (3.6), insertion of these vertex operators is equivalent to the insertion of two-dimensional partitions satisfying interlacing, or transposed interlacing conditions. An infinite sequence of such interlacing partitions effectively builds up a three-dimensional crystal. A relative position of two adjacent slices is determined by a type of two corresponding vertex operators. On the other hand, insertions of color operators have an interpretation of coloring the crystal. The colors 𝑄𝑖 appear in the same order in each composite operator, so these colors are always repeated periodically in the full correlators. Therefore, three-dimensional crystals are built of interlacing, periodically colored slices.

To get more insight about a geometric structure of a crystal, it is convenient to introduce the following graphical representation. We associate various arrows to the vertex operators, as shown in Figure 9. These arrows follow the order of the vertex operators in the fermionic correlators and are drawn from left to right, or up to down (either of these directions is independent of the orientation of the arrow). Following the order of the vertex operators in a given correlator, and drawing a new arrow at the end of the previous one, produces a zig-zag path which represents a shape of the crystal. The coloring of the crystal is taken care of by keeping track of the order of 𝑄𝑖 operators, and by drawing at the endpoint of each arrow a (dashed) line, rotated by 45°, colored according to 𝑄𝑖 which we come across. These lines represent two-dimensional slices in appropriate colors. In this way, the corners of two-dimensional partitions arising from slicing of the crystal are located at the end-points of the arrows. The orientation of arrows represents the interlacing condition (i.e., arrows point from a larger to smaller partition). The interlacing pattern between two consecutive slices corresponds to the types of two consecutive arrows. Finally, the points from which two arrows point outwards represent those stones in the crystal, which can be removed from the initial, full crystal configuration. In fermionic correlators, these points correspond to Γ𝑡𝑖+ followed by Γ𝑡𝑗 operators. We illustrate this graphical construction in a few examples in the next section.

357016.fig.009
Figure 9: Assignment of arrows.
4.4. Examples
4.4.1. Revisiting 3

Let us reconsider 3 geometry which motivated our discussion in Section 4.1. In this case, the dual toric diagram consists just of one triangle, see Figure 10(a), so there is just one vertex and only one color 𝑄0𝑄, and the operators (4.4) take form 𝐴±=Γ±(1)𝑄.(4.37) In consequence, the BPS partition function (4.10) takes exactly the form (4.1).

357016.fig.0010
Figure 10: Toric diagram for 3 (a) consists of one vertex. Operators 𝐴± involve a single Γ± and have a simple arrow (lower part of (a)), as follows from Figure 9. The correlator (4.1) is translated into a sequence of arrows, with rotated dashed lines representing insertions of interlacing two-dimensional partitions. The resulting figure (b) represents plane partitions crystal model, the same as in Figure 7, but now seen from the bottom.

The crystal structure can be read off from a sequence of arrows associated to 𝐴± operators, following the rules in Figure 9. This gives rise to the crystal shown in Figure 10(b). This is the same crystal as in Figure 7, which represents plane partitions, however, now seen from the opposite side.

4.4.2. Orbifolds 3/𝑁+1

Now we consider the resolution of 3/𝑁+1 orbifold. In this case, the toric diagram takes form of a triangle of area (𝑁+1)/2, see Figure 11(a). There are 𝑁 independent 1's and 𝑁+1 vertices of the same 𝑡𝑖=+1, and operators in (4.4) take the form 𝐴±=Γ±𝑄(1)1Γ±𝑄(1)2Γ±𝑄(1)𝑁Γ±𝑄(1)0.(4.38) In the noncommutative chamber, the corresponding crystal consists of plane partitions, however, with slices colored periodically in 𝑁+1 colors. The partition function in the noncommutative chamber is given by (4.10).

fig11
Figure 11: Toric diagram for the resolution of 3/𝑁+1 geometry has 𝑁+1 vertices of the same type . (a): toric diagram for 𝑁=1 and arrow representation of 𝐴+ and 𝑊1. In the noncommutative chamber, this leads to the same plane partition crystal as in Figure 10, however colored now in yellow and red. (b): for the chamber with positive 𝑅 and 2<𝐵<3, the crystal develops two additional corners and its generating function reads 𝑍2|1=Ω+|(𝑊1)2|Ω. (c): for negative 𝑅 and positive 𝑛1<𝐵<𝑛 the crystal is finite along two axes and develops 𝑛1 yellow corners; its generating function for the case of 𝑛=5 shown in the picture reads 𝑍5|1=0|(𝑊1)5|0 (two external arrows, corresponding to Γ acting on 0| and Γ+ acting on |0, are suppressed.).

If we turn on an arbitrary 𝐵-field through a fixed 1, the structure of wall-crossing operators gives rise to modified containers, see for example Figure 11(b). In particular, enlarging the 𝐵-field by one unit adds one more yellow corner to the crystal.

The crystals corresponding to 𝑅<0 are also easy to find. In the extreme chamber, we get a trivial (empty) crystal, representing a single D6-brane (4.13). Adding wall-crossing operators results in a crystal with several corners, finite along two axis (and extending infinitely along the third axis), as shown in Figure 11(c).

4.4.3. Resolved Conifold

We already presented pyramid crystals for the conifold in Section 2.2. They arise from our formalism as follows. The dual toric diagram for the conifold, see Figure 12(a), consists of two triangles and encodes a single (𝑁=1) 1. Two vertices of the toric diagram correspond to two colors 𝑄1 and 𝑄0, so that 𝑄𝑄=1𝑄0,𝑞=𝑞1𝑞0.(4.39) The operators (4.4) in this case read 𝐴±(𝑥)=Γ±(𝑥)𝑄1Γ±(𝑥)𝑄0,(4.40) while (4.5) is𝐴+(𝑥)=Γ+(𝑥𝑞)Γ+𝑥𝑞𝑞1,𝐴(𝑥)=Γ(𝑥)Γ𝑥𝑞1,(4.41) and they satisfy 𝐴+(𝑥)𝐴(𝑦)=1+𝑥𝑦𝑞/𝑞11+𝑥𝑦𝑞𝑞1(1𝑥𝑦𝑞)2𝐴(𝑦)𝐴+(𝑥).(4.42)

fig12
Figure 12: (a): toric diagram for the conifold and arrow representation of 𝐴+ and 𝑊1. (b): for chambers with negative 𝑅 and positive 𝑛1<𝐵<𝑛 the crystals are given by finite pyramid partitions with 𝑛1 additional corners, represented by 𝑛1 stones in the top row (the figure shows the case 𝑛=4). The generating function is given by 𝑍𝑛|1=0|(𝑊1)𝑛|0 which reproduces the result (2.27).

The quantum states (4.9) take form ||Ω=𝐴(1)𝐴(𝑞)𝐴𝑞2||Ω0,+||||=0𝐴+𝑞2𝐴+(𝑞)𝐴+(1),(4.43) and the wall-insertion operators (4.6) are 𝑊1(𝑥)=Γ(𝑥)𝑄1Γ+(𝑥)𝑄0,𝑊1(𝑥)=Γ+(𝑥)𝑄1Γ(𝑥)𝑄0.(4.44)

Therefore, the fermionic correlators take form 𝑍𝑛|1=Ω+||𝑊1𝑛||Ω𝑍,𝑛|1||=0𝑊1𝑛||0.(4.45) and encode generating functions (2.25) and (2.27) introduced in Section 2.2. In the noncommutative chamber, we get the result found first in [14], 𝑍0|1=Ω+Ω, while a single D6-brane is encoded in 𝑍=00=1. These crystals are shown in Figures 12(b) and 13.

357016.fig.0013
Figure 13: Conifold crystal in the chamber with positive 𝑅 and 2<𝐵<3 takes form of pyramid partitions with 3 stones in the top row. Its generating function is given by 𝑍2|1=Ω+|(𝑊1)2|Ω.

5. Matrix Models and Open BPS Generating Functions

In this section, we explain how matrix model formalism can be applied to analyze BPS counting functions. In the first part, Section 5.1, we explain how to relate fermionic formalism, derived in the previous section, to matrix model representation. In Section 5.2, we illustrate how to construct matrix models for the closed noncommutative chamber. In Section 5.3, we analyze in detail BPS generating functions for the conifold for all chambers with 𝑅>0, and derive corresponding spectral curves. We discuss how these curves relate to (and generalize) mirror curves, which we find (as we should) in the commutative chamber. In Section 5.4, we reveal that matrix model representation in fact encodes open BPS generating functions, which can be identified with matrix model integrands.

5.1. Matrix Models from Free Fermions

Let us explain how to relate fermionic representation of BPS amplitudes, introduced in Section 4.2, to matrix models. This relies on introducing into fermionic correlators representing BPS generating functions, such as (4.10) or (4.23), a special representation of the identity operator 𝕀. The representation we are interested in also consists of infinite product of vertex operators and arises as follows [29]. Firstly, we can use the representation as a complete set of states 𝕀=|𝑅𝑅|, which represent two-dimensional partitions. Using orthogonality relations of 𝑈() characters 𝜒𝑅, and the fact that these characters are given in terms of Schur functions 𝜒𝑅=𝑠𝑅(𝑧) for 𝑧=(𝑧1,𝑧2,𝑧3,), we can write 𝕀=𝑅||||=𝑅𝑅𝑃,𝑅𝛿𝑃𝑡𝑅𝑡||||=𝑃𝑅𝒟𝑈𝑃,𝑅𝑠𝑃𝑡𝑧𝑠𝑅𝑡||||=𝑧𝑃𝑅𝒟𝑈𝛼Γ𝑧𝛼||||00𝛼Γ+𝑧𝛼1.(5.1) When such a representation of the identity operator is introduced into (4.10) or (4.23) (or any other correlator of similar structure), we can commute away Γ𝑡𝑖± operators and get rid of operator expressions. For example, inserting the above identity operator in the string of 𝐴+ operators in (4.16) leads to a matrix model with the unitary measure 𝑍𝑛||=0𝑖=𝑘𝐴+||𝕀||(1)𝑘1𝑗=0𝐴+(1)𝑊𝑛||Ω=||𝒟𝑈0𝑖=𝑘+1𝐴+(1)𝛼Γ𝑧𝛼||||00𝛼Γ+𝑧𝛼1𝑘𝑗=0𝐴+(1)𝑊𝑛||Ω=𝑓𝑘𝑛𝑞,𝑄𝑖𝒟𝑈𝛼𝑒(1/𝑔𝑠)𝑉𝑘𝑛(𝑧𝛼).(5.2) The product over 𝛼 represents distinct eigenvalues 𝑧𝛼. Note that we have inserted 𝕀 at the position 𝑘 in the string of 𝐴+(1) operators. In particular, this affects the form of the resulting potential 𝑉𝑘𝑛(𝑧). Moreover, apart from matrix integral, we find some overall factors 𝑓𝑘𝑛(𝑞,𝑄𝑖) which take form of various infinite products. They arise, in a generic chamber, from commutations between Γ± ingredients of wall-crossing operators, and Γ ingredients of |Ω states. In the closed noncommutative chamber 𝑛=0, these factors are trivial, 𝑓𝑛=0(𝑞,𝑄𝑖)=1, and they largely simplify in the commutative chamber 𝑛.

There is a large freedom in choosing the value of k, and it is natural to ask if this choice has some physical interpretation. It was argued in [27] that this is indeed the case, and the choice of 𝑘 is equivalent to the choice of open BPS chamber (open BPS chambers were introduced in Section 2.1). In particular, it turns out that the open generating parameter can be identified with matrix eigenvalues 𝑧𝛼, and the open BPS generating function (2.16) in the open chamber labeled by 𝑘 can be identified with matrix integrand 𝒵openBPS=𝑒(1/𝑔𝑠)𝑉𝑘𝑛(𝑧).(5.3) Even though the overall factors 𝑓𝑘𝑛 in (5.1) may involve closed moduli 𝑄𝑖, they do not involve open moduli 𝑧. In this sense, the matrix integrand is well defined, and up to some simple identification can be identified with open BPS generating function. This identification of parameters amounts to the shift 𝑧𝑧𝑞1/2 (to match earlier M-theory convention with half-integer powers of 𝑞, to integer powers of 𝑞 in the fermionic formalism), as well as identification of Kähler parameters considered in M-theory derivation with parameters 𝜇𝑖 introduced below. We also note that the BPS generating function in (2.16) is determined by the open topological string partition function associated to the external axis of the toric diagram, as in Figure 14. As we will also see, the value of the above integral (5.1) can be related to some more general Calabi-Yau geometry 𝑌.

357016.fig.0014
Figure 14: Brane associated to the external leg of a toric diagram (of a conifold in this particular case). Closed string parameter is denoted by 𝑄 and open string parameter by 𝑧.
5.2. Matrix Models for the Noncommutative Chamber

In this section, we illustrate the relation between BPS counting and matrix models in case of the noncommutative chamber 𝑛=0, and the choice of open chamber also 𝑘=0. This corresponds to the insertion of the identity representation (75) exactly in between |Ω± states in (4.10). In this 𝑛=0 case no factor 𝑓𝑘𝑛 in (5.1) arises, and we obtain matrix models with potentials which can be expressed in terms of the following version of the theta function: Θ(𝑧;𝑞)=𝑗=01+𝑧𝑞𝑗𝑞1+𝑗+1𝑧.(5.4) For a general geometry of the form shown in Figure 8, with types of vertices given by 𝑡𝑖, corresponding matrix models take form 𝑍=𝑑𝑈𝛼𝑒(1/𝑔𝑠)𝑉(𝑧𝛼)=𝑑𝑈𝛼𝑁𝑙=0Θ𝑡𝑙+1𝑧𝛼𝑞1𝑞𝑙;𝑞𝑡𝑙+1,(5.5) where integral is over unitary matrices of infinite size, 𝑁=. Special cases of this result include (i)for 3, the result (5.5) provides a matrix model representation of MacMahon function 𝑍=𝑀(1)=𝑘=1(1𝑞𝑘)𝑘 in terms of a matrix model of the form (5.5) with the integrand 𝑒(1/𝑔𝑠)𝑉(𝑧)=𝑗=01+𝑧𝑞𝑗𝑞1+𝑗+1𝑧=Θ(𝑧;𝑞);(5.6)(ii)for the conifold, we obtain a representation of the pyramid partition generating function (2.23) (with 𝑛=0) in terms of a matrix model with the integrand 𝑒(1/𝑔𝑠)𝑉(𝑧)=𝑗=01+𝑧𝑞𝑗1+𝑞𝑗+1/𝑧(1+𝑄𝑧𝑞𝑗)𝑞1+𝑗+1=/𝑄𝑧Θ(𝑧;𝑞)Θ(𝑄𝑧;𝑞);(5.7)(iii)for 3/𝑁+1, we have 𝑡𝑝=+1 for all 𝑝 and we find matrix model representation of the BPS generating function in terms of a matrix model with the integrand 𝑒(1/𝑔𝑠)𝑉(𝑧)=𝑗=01+𝑧𝑞𝑗𝑞1+𝑗+1𝑧𝑞1+1𝑞𝑁𝑧𝑞𝑗𝑞1+𝑗+1𝑞1𝑞𝑁𝑧=𝑁𝑙=0Θ𝑞1𝑞𝑙.𝑧;𝑞(5.8)

5.3. Matrix Model for the Conifold Analysis

In this section, we illustrate how matrix model techniques can be used in the context of models which arise for BPS counting. We focus on the conifold matrix model in arbitrary closed BPS chamber 𝑛, and fixed 𝑘=0. In this case, the result (5.2) takes form (after the redefinition 𝑄=𝑞1𝑞𝑛) 𝑍𝑛=𝑀(𝑞)2𝑗=11𝑄𝑞𝑗𝑗1𝑄1𝑞𝑗+𝑛𝑗+𝑛=𝑓𝑛(𝑞,𝑄)𝑑𝑈𝛼𝑗=01+𝑧𝛼𝑞𝑗+11+𝑞𝑗/𝑧𝛼1+𝑧𝛼𝑞𝑗+𝑛+1/𝑄1+𝑞𝑗𝑄/𝑧𝛼=𝑓𝑛(𝑞,𝑄)𝑍matrix,(5.9) with 𝑓𝑛(𝑞,𝑄)=𝑀(𝑞)𝑗=11𝑞𝑛+𝑗/𝑄𝑛𝑀(𝑞𝑛,𝑞),(5.10) with MacMahon function 𝑀(𝑞) defined in (2.17), and with the following generalized MacMahon function 𝑀(𝑧,𝑞)=𝑖=11(1𝑧𝑞𝑖)𝑖.(5.11) In particular, in the noncommutative chamber 𝑓conifold0=1, and in the commutative chamber 𝑓conifold𝑛=𝑀(𝑞) which represents topological string degree zero contributions. The result (5.9) implies that the value of the matrix model integral (without the prefactor 𝑓𝑛) is equal to 𝑍matrix=𝑍𝑛𝑓𝑛(𝑞,𝑄)=𝑀(𝑞)𝑗=11𝑄𝑞𝑗𝑗1𝜇𝑞𝑗𝑗(1𝜇𝑄𝑞𝑗)𝑗=𝑑𝑈𝛼𝑗=01+𝑧𝛼𝑞𝑗+11+𝑞𝑗/𝑧𝛼1+𝑧𝛼𝑞𝑗+𝑛+1/𝑄1+𝑞𝑗𝑄/𝑧𝛼,(5.12) where 𝜇=𝑞𝑛/𝑄.

Now we wish to analyze the matrix model 𝑍matrix. We parametrize the 't Hooft coupling and the chamber dependence, respectively, by 𝑇=𝑔𝑠𝑁,𝜏=𝑛𝑔𝑠.(5.13) As our models correspond to 𝑈() matrices, ultimately we are interested in the limit 𝑇,𝑔𝑠=const,𝑄=const,(5.14) for each fixed chamber (i.e., fixed 𝑛 and therefore 𝜏). The noncommutative chamber corresponds to 𝜏=0, while 𝜏 represents the topological string chamber.

Using the expansion of the quantum dilogarithm log𝑖=11𝑧𝑞𝑖1=𝑔𝑠𝑚=0Li2𝑚𝐵(𝑧)𝑚𝑔𝑚𝑠𝑚!,(5.15) and the redefinition of the unitary measure (3.17) we find, to the leading order in 𝑔𝑠, the following matrix model potential: 𝑉𝜏=𝑇log(𝑧)+Li2(𝑧)+Li21𝑧Li2𝑄𝑧Li2𝑧𝑄𝑒𝜏,(5.16) so that 𝜕𝑧𝑉𝜏=𝑇log(𝑧+𝑄)+log(1+(𝑧/𝑄𝑒𝜏))𝑧.(5.17)

Now we wish to solve the model (5.9) in the small 𝑔𝑠 limit. Firstly, we need to find the resolvent 𝜔(𝑝), which can be done using the Migdal integral (3.18), and careful derivation is presented in [29]. As we expect one-cut solution of our model, from the Migdal integral we get an expression in terms of the end-points of this cut 𝑎 and 𝑏. The normalization condition (3.19) imposes two constraints, for terms of order 𝑝0 and 𝑝1 in the resolvent, which take form 𝑎+𝑄𝑏+𝑄𝑎+𝑄𝑒𝜏𝑏+𝑄𝑒𝜏=𝑄1/2𝑒(𝜏+𝑇)/2,(𝑎+𝑄)𝑏(𝑏+𝑄)𝑎(𝑎+𝑄𝑒𝜏)𝑏(𝑏+𝑄𝑒𝜏)𝑎=𝑄1/2𝑒(𝜏+𝑇)/2.(5.18) These constraints can be solved in the exact form, with result 𝑎=1+𝜖2(1𝜇)1𝜇𝜖2+(1𝑄)1+𝜇𝜖22𝜇1𝜇𝜖22+2𝑖𝜖(1𝑄)1𝜖2(1𝜇)1𝑄𝜇𝜖21𝜇𝜖22,𝑏=1+𝜖2(1𝜇)1𝜇𝜖2+(1𝑄)1+𝜇𝜖22𝜇1𝜇𝜖222𝑖𝜖(1𝑄)1𝜖2(1𝜇)1𝑄𝜇𝜖21𝜇𝜖22,(5.19) where we introduced 𝜖=𝑒𝑇/21,𝜇=𝑄𝑒𝜏.(5.20) Substituting these end-points back to the formula for the resolvent, we find 𝜔±1(𝑝)=𝑝𝑇log1+𝜇𝜖2𝑝+1+𝑄𝜖21𝜇𝜖2(𝑝𝑎)(𝑝𝑏)2𝑒𝑇(𝑝+𝑄).(5.21) As a check, this result indeed satisfies the consistency condition (3.20) 𝜔+(𝑝)+𝜔𝜕(𝑝)=𝑝𝑉𝜏(𝑝)𝑇,(5.22) with 𝑉𝜏 given in (5.16). From the knowledge of the resolvent, we can also determine eigenvalue density along the cut (3.21) 1𝜌(𝑝)=𝑝𝑇log1+𝜇𝜖2𝑝+1+𝑄𝜖21𝜇𝜖2(𝑝𝑎)(𝑝𝑏)1+𝜇𝜖2𝑝+1+𝑄𝜖2+1𝜇𝜖2(𝑝𝑎)(𝑝𝑏),(5.23) as well as the spectral curve. Writing 𝑥=𝑝𝑇𝜔(𝑝) and 𝑝=𝑒𝑦, and after a few simple rescalings, we find that the spectral curve takes form 𝑒𝑥+𝑦+𝑒𝑥+𝑒𝑦+𝑄1𝑒2𝑥+𝑄2𝑒2𝑦+𝑄3=0,(5.24) where 𝑄1=𝜖21+𝜇𝑄1+𝜇𝜖21+𝑄𝜖2,𝑄2=𝜇1+𝑄𝜖2(1+𝜇𝑄)1+𝜇𝜖2,𝑄3=𝑄1+𝜇𝜖21+𝜖2𝑄.(1+𝜇𝑄)(5.25)

The above curve is given by a symmetric function of 𝑄, 𝜇=𝑄1𝑞𝑛 and 𝜖2=𝑒𝑇. Apparently this is a mirror curve of the so-called closed topological vertex geometry, which is a Calabi-Yau manifold arising from a symmetric resolution of 3/2×2 orbifold, see Figure 15. This geometry has three moduli, that is, the original Kähler moduli 𝑄 of the resolved conifold, the chamber parameter 𝑛 (encoded in 𝜇), and the 't Hooft parameter 𝑇, which are all unified in a geometric way in our matrix model. Moreover, the fractional coefficients in the curve equation (5.24) encode the correct mirror map for the closed topological vertex geometry (and to the linear order, these coefficients are just 𝑄, 𝜇, and 𝑒𝑇).

357016.fig.0015
Figure 15: The spectral curve for the conifold matrix model (5.9) in arbitrary closed BPS chamber coincides with the mirror curve of the closed topological vertex geometry, whose toric diagram is shown above. This geometry has three 1 moduli, which encode conifold size 𝑄, the closed BPS chamber via 𝜇, and finite 't Hooft coupling via 𝑒𝑇.

In the BPS counting problem we are interested in, as follows from the form of the identity operator (5.1), ultimately we need to consider matrices of infinite size. We also need to keep fixed 𝑔𝑠, so we should consider the limit of 𝑇, or equivalently 𝜖0. Up to a linear shift of 𝑥 and 𝑦, the (5.24) in this limit becomes 𝜇𝑒2𝑦+𝑒𝑥+𝑦+𝑒𝑥+(1+𝑄𝜇)𝑒𝑦+𝑄=0.(5.26) The manifold corresponding to this curve is the suspended pinch point (SPP) geometry, with 𝑄 and 𝜇 encoding flat coordinates representing sizes of its two 1's. Having found the SPP mirror curve, let us also make the following remarks.

Firstly, we see that not only the spectral and mirror curves agree, but moreover the matrix integral (5.9) reproduces (after the identification 𝑞=𝑞𝑠) the full topological string partition function of the SPP geometry at finite 𝑔𝑠, which is known to take form 𝒵SPPtop𝑞𝑠=,𝑄,𝜇𝑘=11𝑄𝑞𝑘𝑠𝑘1𝜇𝑞𝑘𝑠𝑘1𝑞𝑘𝑠3𝑘/21𝜇𝑄𝑞𝑘𝑠𝑘,(5.27) for Kähler parameters 𝑄 and 𝜇. This confirms that our result makes sense, although this also means that the terms in lowest order in 𝑔𝑠 in the potential reproduce the full 𝑔𝑠 dependence of the partition function. It would be nice to prove rigorously that higher 𝑔𝑠 corrections to the potential indeed do not affect the total partition function. This appears to be a very special feature of matrix models integrands which can be expressed—as is the case for (5.12)—in terms of infinite products of the form 𝑘(1𝑥𝑞𝑘). One proof of such statement in a related situation (although in addition taking advantage of a special phenomenon of the arctic circle) can be found in [52].

Secondly, it is natural to conjecture that the total partition function of the matrix model, for finite 𝑇, should reproduce (at least up to some MacMahon factor) the topological string partition function for the closed topological vertex which reads 𝑍totalmatrix𝑞,𝑄,𝜇,𝜖2=𝑘=11𝑞𝑘𝑘𝑘=11𝑄𝑞𝑘𝑘1𝜇𝑞𝑘𝑘1𝜖2𝑞𝑘𝑘1𝑄𝜇𝜖2𝑞𝑘𝑘1𝑄𝜇𝑞𝑘𝑘1𝜇𝜖2𝑞𝑘𝑘1𝑄𝜖2𝑞𝑘𝑘.(5.28)

Finally, we also note that in the limit 𝑄,𝜇0 our model reduces to the Chern-Simons matrix model discussed in [47, 50]. Indeed, in this limit the potential (5.16) reproduces Gaussian potential [50] 𝑉𝑄0,𝑛1=2(log𝑧)21=2𝑢2,(5.29) and the partition function (5.9) correctly reduces to the appropriate Chern-Simons partition function. In this case, the resolvent (5.21) reduces to 𝜔conifold±(𝑝)𝜇=𝑄=0=1