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Integrable fishnet circuits and Brownian solitons
by Žiga Krajnik, Enej Ilievski, Tomaž Prosen, Benjamin J. A. Héry, Vincent Pasquier
This is not the latest submitted version.
Submission summary
Authors (as registered SciPost users): | Žiga Krajnik |
Submission information | |
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Preprint Link: | https://arxiv.org/abs/2411.08030v1 (pdf) |
Date submitted: | 2024-11-14 04:21 |
Submitted by: | Krajnik, Žiga |
Submitted to: | SciPost Physics |
Ontological classification | |
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Academic field: | Physics |
Specialties: |
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Approaches: | Theoretical, Computational |
Abstract
We introduce classical many-body dynamics on a one-dimensional lattice comprising local two-body maps arranged on discrete space-time mesh that serve as discretizations of Hamiltonian dynamics with arbitrarily time-varying coupling constants. Time evolution is generated by passing an auxiliary degree of freedom along the lattice, resulting in a `fishnet' circuit structure. We construct integrable circuits consisting of Yang-Baxter maps and demonstrate their general properties, using the Toda and anisotropic Landau-Lifschitz models as examples. Upon stochastically rescaling time, the dynamics is dominated by fluctuations and we observe solitons undergoing Brownian motion.
Author indications on fulfilling journal expectations
- Provide a novel and synergetic link between different research areas.
- Open a new pathway in an existing or a new research direction, with clear potential for multi-pronged follow-up work
- Detail a groundbreaking theoretical/experimental/computational discovery
- Present a breakthrough on a previously-identified and long-standing research stumbling block
Current status:
Reports on this Submission
Strengths
The manuscript contains novel ideas for the time discretization of the integrable dynamics
Weaknesses
See report
Report
The manuscript introduces a new approach to the space and time discretization dynamics of integrable systems. The construction of the chiral propagator seems novel and exciting, even though the connected two-body map has appeared before (e.g. Eq. 4.46 in [11]).
I have conceptual problems with understanding this work and some minor comments.
1. I wonder what is the connection between the time dynamics generated by the chiral operator and any dynamics generated by the Poisson structure.
More specifically, is it possible to interpret Eq. (3.6) as a group version of $\{O,H\}$ with some H?
As the paper is written now, the logic is a bit broken: in (2.22) the commuting flows are introduced, however, the time dynamics is not generated by them!
2. Perhaps a related question is the canonicity of the chiral operator.
I am confused by the discussion at the bottom of page 11.
The Poisson structure is given by the r-matrix (2.1) and is not necessarily canonical (for example the algebra of spins (4.35)).
This suggests that any proof of canonicity should be r-matrix-specific.
Surely, locally, in the specific coordinates the Poisson structure is given by (2.35) but do we know the chiral operator in these coordinates?
Additionally, for the Toda chain example, where the Poisson brackets are given by (2.35), the two-body map is NOT canonical!
On top of everything, the following sentence after Eq. (4.15) is very confusing.
" A straightforward calculation shows that the map $\psi_\tau$ is not canonical. Nevertheless, the chiral propagators $\Phi_\tau^\pm$ are canonical, as already established by Sklyanin [23,24,35]."
So, is it actually that the chiral propagators have been introduced by Sklyanin and not by the authors? What part of the proof in Appendix A fails for Toda's two-body map?
3. From the introduction, it seems that the desired approach to the stochastic integrable systems is to keep integrable dynamics and introduce randomness in the initial conditions. However, the approach of Sec. 5 is different; the dynamics is replaced by the non-integrable averaged propagator (5.5) (as the charges (3.7) cannot be conserved under the linear combination of the chiral propagators unless they are the linear functions). I think discussion of this point should be beneficial for the readers. Additionally, beyond trajectory analysis, examining the time dependence of energy could help to distinguish various regimes.
=======================================================
Minor Comments and Typos
Most minor points and typographical errors have already been addressed in two reports published while this report was in preparation. I will, however, highlight some additional issues and apologize for any redundancies:
1. What does the sign $\simeq$ in Eq. (2.39) mean? is this that the matrices are proportional to identity? Why then the Toda matrices (4.5) and (4.6) do not satisfy this condition?
2. The caption for Fig. 3 references Eq. (2.39), but it might be more appropriate to cite Eq. (1.4) instead.
3. The continuous-time limit for Toda is confusing. Why $\tau\to\infty$? Why the double-step propagator?
4. In the introduction it is claimed that
"An important step forward in this direction has been the development of generalized hydrodynamics, a hydrodynamic theory for integrable systems [8, 9]. However, at present the assumptions of the theory limit its applicability to leading-order dynamics on the ballistic scale. To access dynamics on sub-ballistic scales a recourse to numerical simulations remains a necessity."
Surely, the GHD on the diffusive level has been studied extensively see, for example [1812.00767].
Once the authors have addressed these issues, I will be happy to recommend this manuscript for publication in SciPost.
Recommendation
Ask for major revision
Report #2 by Vincent Caudrelier (Referee 2) on 2025-1-30 (Invited Report)
Strengths
Please see my report for more details but overall strengths are:
- New exciting results taking known methods and notions in a new direction
- Great potential to bring together various communities and initiate further research.
Weaknesses
Again, more detail in my report but overall (minor) weaknesses are:
- A limited rendition of how the work fits in the context of discrete integrable systems. This is not a judgement and not a problem (of course everyone knows certain areas better than others) and I have tried to make suggestions to improve the impact this work can have in the community of discrete integrable systems, at least.
- Sometimes notations are a bit confusing. Again, I've tried to point out where this happens.
Report
Yes, this article definitely deserves publication in SciPost. It easily fulfills several criteria.
Requested changes
Please see the attached report.
Recommendation
Publish (surpasses expectations and criteria for this Journal; among top 10%)
Author: Žiga Krajnik on 2025-03-12 [id 5287]
(in reply to Report 2 by Vincent Caudrelier on 2025-01-30)
We thank the referee for their detailed reading of the manuscript and numerous useful comments which have helped to improved the manuscript. We address their points below:
- We thank the referee for point out this typo which has been corrected.
- As noted by the referee, the literature on discrete and semi-discrete integrable systems is vast and we do not feel qualified to give an overview, however partial. We therefore thank the referee for pointing out these references which have been incorporated into the manuscript.
- As noted, the factorization Eq. 1.1 (equivalently Eq. 2.2) is the basic building block for the outlined construction. One key feature that might have been missed by the referee is that while Eq. 2.2 allows for two free parameters, $\lambda$ and $\tau$, we eventually only consider solutions that depend on a single parameter $\tau$, see Eq. 2.3. In other words, we consider solutions that are of difference form: we therefore deal with a single complex spectral parameter $\lambda$, and treat $\tau$ as a fixed lattice parameter. Hopefully this also clarifies several other points raised by the referee. We thank the referee for bringing the reference on entwining Yang-Baxter maps to our attention, which has been incorporated into the manuscript after Eq. 2.3.
- We now explicitly mention that the transfer matrices $\mathcal{T}^{(n)}$ are obtained by tracing powers of the monodromy matrix before Eq. 1.5.
- As we now state explicitly after Eq. 2.1, by `mutual compatibility' we mean that both Lax operators obey the quadratic Sklyanin bracket with the same $r$-matrix. We also note that $\mathcal{L}$ and $\mathcal{L}^a$ are defined on different, in general non-isomorphic, phase-spaces and therefore their matrix elements trivially Poisson-commute.
- In 2.8 we use the refactorization $\mathcal{L}^a_{\lambda^-}(y^\prime)$ $\mathcal{L}_{\lambda^+-\xi}(x^\prime)$ $ = \mathcal{L}_{\lambda^+-\xi} (x)$ $\mathcal{L}^a_{\lambda^-}(y)$. Note that only $\mathcal{L}$ is shifted by the inhomogeneity. As already noted in our response to the third query, we specialize to maps $\psi$ that are of difference form. We can therefore absorb the inhomogeneity $\xi_\ell$ by (locally) shifting the time-step parameter as $\tau \to \tau - \xi_\ell$ yielding $\psi_{\tau - \xi_\ell}$ where $\psi$ is defined as the solution of 2.2. We therefore find it appropriate to say that Eq. 2.7 is resolved by the repeated application of Eq. 2.2. If both Lax operators were shifted by the inhomogeneity all dependence on $\xi_\ell$ could be eliminated by a shift of the spectral parameter $\lambda \to \lambda - \xi_\ell$. The reasons for introducing inhomogeneities in the manuscript was to exhibit the most general form of the construction. In the language of vertex models this means that all parameters on vertical (inhomogeneities) and horizontal (time-step parameters) lines can be chosen independently. The reason for specializing to space-homogeneous models in our applications was to ensure parity of the resulting dynamical systems. While we could also recover parity at the level of inhomogeneity-averaged observables by their appropriate sampling, we have chosen to focus on time-inhomogeneity so as not to unnecessarily obscure the message. We believe the study of space-time inhomogeneous models is an interesting direction for further study.
- We thank the referee for catching this typo which has been corrected.
- We thank the referee for point this out. We have corrected Eq. 2.21 so that it also holds for $n,m \geq 1$.
- As we note after Eq. 2.25, the expansion in Eq. 2.23 merely generates conserved quantities without regard to their locality properties. To generate local conserved quantities one needs to expand the transfer functions around a distinguished (but model-specific) point $\lambda_0$ at which the Lax operators degenerate. In the particular case of the Toda chain local conserved quantities arise by expanding around $\lambda \to \infty$.
- While taking the trace of Eq. 2.7 does indeed produce Eq. 2.19 for transfer matrices with spectral parameter $\lambda^+$, one should keep in mind that $\lambda$ is a free parameter. The reported relation 2.19 is obtained by shifting $\lambda \to \lambda - \tau/2$, as we now comment explicitly above Eq. 2.19.
- We have reorganized the discussion of the Yang-Baxter relation, and now we start by introducing the intertwiner. As we comment after Eq. 2.27 an additional assumption of uniqueness of factorization is necessary to obtain the Yang-Baxter equation from the equivalence of the products of Lax operators, see also footnote 4. In our reading, this is the content of Proposition 3.1 of reference [3].
- To understand the passage from Eq. 2.33 to 2.34 consider Figure 6 (starting on the right-hand side): the physical variables $X$ which are injected at the bottom are subsequently sequentially propagated by the propagators defined by the fixed points $y_2^*$ and $y_1^*$, resulting in the updated physical variables $X'$ exiting on top, achieved by a composition of two left propagators. The pair of fixed points is then mapped by the intertwiner to ${y_1^*}'$ and ${y_2^*}'$, respectively. Note that at this point we have not assumed any particular properties of either ${y_2^*}'$ or ${y_1^*}'$. Consider next the left-hand side of Figure 6. We again input physical variables $X$ at the bottom and $y_2^*$ and $y_1^*$ on the right. Before propagating $X$, the pair of auxiliary variables is first mapped by the intertwiner to ${y_1^*}'$ and ${y_2^*}'$ as defined on the right-hand side. By considering the output auxiliary variables on the right-hand side of Figure 6, we now observe that ${y_1^*}'$ and ${y_2^*}'$ are in fact fixed points of the maps $\Psi_{\tau'}$ and $\Psi_{\tau}$ on the left-hand side, respectively. The propagation of $X$ to $X'$ on the left-hand side is therefore also achieved by the composition of two left propagators. To recognize the significance of the assumption, namely that the intertwiner does not mix fixed points, we emphasize that the propagators are (in case of multiple fixed points) implicitly defined by the choice of fixed point as we note after Eq. 2.18. To obtain the commutation of propagators in Eq. 2.34 from Figure 6 (equivalently Eq. 2.33) it is therefore necessary to assume that the intertwiner maps fixed points of the same type into each other. To give an example, consider the Landau-Lifschitz model where there is a pair of fixed points, an attractive and a repulsive one. The assumption prevents the possibility that a pair of attractive fixed points is mapped to a pair of repulsive fixed points or vice-versa. While we believe that this holds on general grounds, we unfortunately do not see how to explicitly demonstrate this at present. We have expanded the footnote to better clarify the meaning of our assumption.
- We now explicitly define $\sigma^0$ after Eq. 4.7 to avoid confusion.
- We are grateful to the referee for spotting this mistake. We have rewritten Eqs. 4.8-4.23 with their comments in mind and corrected the mistake in Eq. 4.17. We hope the improved version is now more easily understandable.
- One of the motivations for our work was the construction of discretizations that preserve integrability of the underlying continuous-time dynamics. We accordingly felt that that identifying the continuous-time limit naturally belongs to the story, especially since it turned out that the limiting procedure for the Toda lattice is not completely straightforward, involving sending $\tau \to \infty$ and considering the composition of a pair of propagators.
- We now use a different font to denote the homology groups.
- To aid the identification of Eq. 4.44 with Figure 2 we now give the explicit identification of variables above Eq. 4.44. The variables, $\zeta^a$ stood for the anisotropic stereographic variable of the auxiliary spin, as noted below Eq. 4.69. To avoid any possible confusion over the superscript $a$ we have now changed $\zeta^a \to \zeta^{\rm aux}$. We also specify $X$ in terms of $\zeta$ below Eq. 4.71 and write out the constant $\alpha_\ell$ explicitly in terms of the matrix elements of the Lax operator. 2.We have modified the notation in the discussion on the convergence of fixed point iteration, see Eqs. 4.81-4.89, with the aim of making it more easily understandable. The crucial observation is that, in the limit of a long chain, the average of the sum is given by a uniformly distributed physical spin. The computation is facilitated by noting that the logarithm decouples the integral over the product of factors involving the two auxiliary spins into the sum of two identical integrals which are easily computed.
- We thank the referee for catching this typo which has been corrected.
While we are not experts on periodic reductions mentioned by the referee, our understanding is that, while similar to our construction, there are several discernible differences. Periodic reductions apply to multilinear quadrilateral maps defined implicitly by $Q(u, \tilde u, \hat u, \tilde{ \hat u}) = 0$ which is solved by '$3\to 1$' maps $\hat u = F(u, \tilde u, \tilde{ \hat u})$. The initial conditions are given on an initial saw-tooth in the space direction and propagated by alternating application of the '$3 \to 1$' map. Periodic closure of the saw-tooth gives a pair of additional constraints which allows for the elimination of two variables.
By contrast, we deal with '$2\to 2$' maps and specify the physical initial condition along a straight line. Time evolution is then defined at expense of introducing an extra initial auxiliary variable which is sequentially propagated horizontally across the system by sequentially scattering with all of the physical variables. Imposing the periodic closure in the auxiliary direction gives rise to a fixed point condition whose solution permits to eliminate the auxiliary variables from the propagator, thereby defining the time-evolution purely in terms of the physical variables. Our construction therefore effectively eliminates half of the involved variables, unlike the periodic reductions which eliminate only two variables.
Brownian motion of the solitons arises due to a local stochastic rescaling of time generated by multiplicative noise. The increments of a soliton's position is proportional to the time-step. As these are independent and (in the continuous-time limit) normally distributed so are the soliton's increments, a characteristic of Brownian motion. Importantly, it is shown that this does not break integrability of the dynamics since the propagators conserve the transfer matrix for all time-steps, unlike an additive noise in e.g. a Langevin equation which would immediately break all conservation laws.
Author: Žiga Krajnik on 2025-03-12 [id 5286]
(in reply to Report 2 by Vincent Caudrelier on 2025-01-30)
We thank the referee for their detailed reading of the manuscript and numerous useful comments which have helped to improved the manuscript. We address their points below:
- We thank the referee for point out this typo which has been corrected.
- As noted by the referee, the literature on discrete and semi-discrete integrable systems is vast and we do not feel qualified to give an overview, however partial. We therefore thank the referee for pointing out these references which have been incorporated into the manuscript.
- As noted, the factorization Eq. 1.1 (equivalently Eq. 2.2) is the basic building block for the outlined construction. One key feature that might have been missed by the referee is that while Eq. 2.2 allows for two free parameters, $\lambda$ and $\tau$, we eventually only consider solutions that depend on a single parameter $\tau$, see Eq. 2.3. In other words, we consider solutions that are of difference form: we therefore deal with a single complex spectral parameter $\lambda$, and treat $\tau$ as a fixed lattice parameter. Hopefully this also clarifies several other points raised by the referee. We thank the referee for bringing the reference on entwining Yang-Baxter maps to our attention, which has been incorporated into the manuscript after Eq. 2.3.
- We now explicitly mention that the transfer matrices $\mathcal{T}^{(n)}$ are obtained by tracing powers of the monodromy matrix before Eq. 1.5.
- As we now state explicitly after Eq. 2.1, by `mutual compatibility' we mean that both Lax operators obey the quadratic Sklyanin bracket with the same $r$-matrix. We also note that $\mathcal{L}$ and $\mathcal{L}^a$ are defined on different, in general non-isomorphic, phase-spaces and therefore their matrix elements trivially Poisson-commute.
- In 2.8 we use the refactorization $\mathcal{L}^a_{\lambda^-}(y^\prime)$ $\mathcal{L}_{\lambda^+-\xi}(x^\prime)$ $ = \mathcal{L}_{\lambda^+-\xi} (x)$ $\mathcal{L}^a_{\lambda^-}(y)$. Note that only $\mathcal{L}$ is shifted by the inhomogeneity. As already noted in our response to the third query, we specialize to maps $\psi$ that are of difference form. We can therefore absorb the inhomogeneity $\xi_\ell$ by (locally) shifting the time-step parameter as $\tau \to \tau - \xi_\ell$ yielding $\psi_{\tau - \xi_\ell}$ where $\psi$ is defined as the solution of 2.2. We therefore find it appropriate to say that Eq. 2.7 is resolved by the repeated application of Eq. 2.2. If both Lax operators were shifted by the inhomogeneity all dependence on $\xi_\ell$ could be eliminated by a shift of the spectral parameter $\lambda \to \lambda - \xi_\ell$. The reasons for introducing inhomogeneities in the manuscript was to exhibit the most general form of the construction. In the language of vertex models this means that all parameters on vertical (inhomogeneities) and horizontal (time-step parameters) lines can be chosen independently. The reason for specializing to space-homogeneous models in our applications was to ensure parity of the resulting dynamical systems. While we could also recover parity at the level of inhomogeneity-averaged observables by their appropriate sampling, we have chosen to focus on time-inhomogeneity so as not to unnecessarily obscure the message. We believe the study of space-time inhomogeneous models is an interesting direction for further study.
- We thank the referee for catching this typo which has been corrected.
- We thank the referee for point this out. We have corrected Eq. 2.21 so that it also holds for $n,m \geq 1$.
- As we note after Eq. 2.25, the expansion in Eq. 2.23 merely generates conserved quantities without regard to their locality properties. To generate local conserved quantities one needs to expand the transfer functions around a distinguished (but model-specific) point $\lambda_0$ at which the Lax operators degenerate. In the particular case of the Toda chain local conserved quantities arise by expanding around $\lambda \to \infty$.
- While taking the trace of Eq. 2.7 does indeed produce Eq. 2.19 for transfer matrices with spectral parameter $\lambda^+$, one should keep in mind that $\lambda$ is a free parameter. The reported relation 2.19 is obtained by shifting $\lambda \to \lambda - \tau/2$, as we now comment explicitly above Eq. 2.19.
- We have reorganized the discussion of the Yang-Baxter relation, and now we start by introducing the intertwiner. As we comment after Eq. 2.27 an additional assumption of uniqueness of factorization is necessary to obtain the Yang-Baxter equation from the equivalence of the products of Lax operators, see also footnote 4. In our reading, this is the content of Proposition 3.1 of reference [3].
- To understand the passage from Eq. 2.33 to 2.34 consider Figure 6 (starting on the right-hand side): the physical variables $X$ which are injected at the bottom are subsequently sequentially propagated by the propagators defined by the fixed points $y_2^*$ and $y_1^*$, resulting in the updated physical variables $X'$ exiting on top, achieved by a composition of two left propagators. The pair of fixed points is then mapped by the intertwiner to ${y_1^*}'$ and ${y_2^*}'$, respectively. Note that at this point we have not assumed any particular properties of either ${y_2^*}'$ or ${y_1^*}'$. Consider next the left-hand side of Figure 6. We again input physical variables $X$ at the bottom and $y_2^*$ and $y_1^*$ on the right. Before propagating $X$, the pair of auxiliary variables is first mapped by the intertwiner to ${y_1^*}'$ and ${y_2^*}'$ as defined on the right-hand side. By considering the output auxiliary variables on the right-hand side of Figure 6, we now observe that ${y_1^*}'$ and ${y_2^*}'$ are in fact fixed points of the maps $\Psi_{\tau'}$ and $\Psi_{\tau}$ on the left-hand side, respectively. The propagation of $X$ to $X'$ on the left-hand side is therefore also achieved by the composition of two left propagators. To recognize the significance of the assumption, namely that the intertwiner does not mix fixed points, we emphasize that the propagators are (in case of multiple fixed points) implicitly defined by the choice of fixed point as we note after Eq. 2.18. To obtain the commutation of propagators in Eq. 2.34 from Figure 6 (equivalently Eq. 2.33) it is therefore necessary to assume that the intertwiner maps fixed points of the same type into each other. To give an example, consider the Landau-Lifschitz model where there is a pair of fixed points, an attractive and a repulsive one. The assumption prevents the possibility that a pair of attractive fixed points is mapped to a pair of repulsive fixed points or vice-versa. While we believe that this holds on general grounds, we unfortunately do not see how to explicitly demonstrate this at present. We have expanded the footnote to better clarify the meaning of our assumption.
- We now explicitly define $\sigma^0$ after Eq. 4.7 to avoid confusion.
- We are grateful to the referee for spotting this mistake. We have rewritten Eqs. 4.8-4.23 with their comments in mind and corrected the mistake in Eq. 4.17. We hope the improved version is now more easily understandable.
- One of the motivations for our work was the construction of discretizations that preserve integrability of the underlying continuous-time dynamics. We accordingly felt that that identifying the continuous-time limit naturally belongs to the story, especially since it turned out that the limiting procedure for the Toda lattice is not completely straightforward, involving sending $\tau \to \infty$ and considering the composition of a pair of propagators.
- We now use a different font to denote the homology groups.
- To aid the identification of Eq. 4.44 with Figure 2 we now give the explicit identification of variables above Eq. 4.44. The variables, $\zeta^a$ stood for the anisotropic stereographic variable of the auxiliary spin, as noted below Eq. 4.69. To avoid any possible confusion over the superscript $a$ we have now changed $\zeta^a \to \zeta^{\rm aux}$. We also specify $X$ in terms of $\zeta$ below Eq. 4.71 and write out the constant $\alpha_\ell$ explicitly in terms of the matrix elements of the Lax operator. 2.We have modified the notation in the discussion on the convergence of fixed point iteration, see Eqs. 4.81-4.89, with the aim of making it more easily understandable. The crucial observation is that, in the limit of a long chain, the average of the sum is given by a uniformly distributed physical spin. The computation is facilitated by noting that the logarithm decouples the integral over the product of factors involving the two auxiliary spins into the sum of two identical integrals which are easily computed.
- We thank the referee for catching this typo which has been corrected.
While we are not experts on periodic reductions mentioned by the referee, our understanding is that, while similar to our construction, there are several discernible differences. Periodic reductions apply to multilinear quadrilateral maps defined implicitly by $Q(u, \tilde u, \hat u, \tilde{ \hat u}) = 0$ which is solved by '$3\to 1$' maps $\hat u = F(u, \tilde u, \tilde{ \hat u})$. The initial conditions are given on an initial saw-tooth in the space direction and propagated by alternating application of the '$3 \to 1$' map. Periodic closure of the saw-tooth gives a pair of additional constraints which allows for the elimination of two variables.
By contrast, we deal with '$2\to 2$' maps and specify the physical initial condition along a straight line. Time evolution is then defined at expense of introducing an extra initial auxiliary variable which is sequentially propagated horizontally across the system by sequentially scattering with all of the physical variables. Imposing the periodic closure in the auxiliary direction gives rise to a fixed point condition whose solution permits to eliminate the auxiliary variables from the propagator, thereby defining the time-evolution purely in terms of the physical variables. Our construction therefore effectively eliminates half of the involved variables, unlike the periodic reductions which eliminate only two variables.
Brownian motion of the solitons arises due to a local stochastic rescaling of time generated by multiplicative noise. The increments of a soliton's position is proportional to the time-step. As these are independent and (in the continuous-time limit) normally distributed so are the soliton's increments, a characteristic of Brownian motion. Importantly, it is shown that this does not break integrability of the dynamics since the propagators conserve the transfer matrix for all time-steps, unlike an additive noise in e.g. a Langevin equation which would immediately break all conservation laws.
Report #1 by Yuan Miao (Referee 1) on 2025-1-29 (Invited Report)
Strengths
1-- The results are novel and important.
2-- The results can potentially lead to further investigations in the future.
3-- Both analytic and numeric results are solid.
Report
The authors studied a new type of discrete space-time evolution of classical integrable models, dubbed "integrable fishnet circuits". The integrability of the circuits is guaranteed by the Yang-Baxter map, and the authors demonstrated the construction in the examples of Toda lattice and anisotropic Landau-Lifshitz magnet. Due to the choice of the "time parameter" $\tau$, classical solitons that obey Brownian motion are discovered.
The results of the article are new and important, and it will lead to further investigations in various directions, such as the quantum counterparts and generalizations to other models related to higher-rank Lie algebras, etc. I believe that the article deserves to be published in SciPost Physics, once the comments below are properly addressed.
1-- The title "integrable fishnet circuits" needs to be clarified. In the field of integrable models, there exists already a terminology "Fishing-net diagram" (later popularized as "fishnet diagram"), a model of Feynman diagrams of fishnet type that can be solved by an integrable vertex model [R1]. Recently, more studies such as [R2] and [R3] have been performed in this field, which has a close relation to AdS/CFT integrability. Even though the content of the current article is not directly related to the research mentioned above, it is worth pointing out the existing literatures on those investigations and the difference/connection to current article, in order to avoid confusions to people who might be familiar with both topics.
2-- As mentioned in Sec. 1.1, the construction of fishnet circuits resembles the quantum transfer matrix formalism, can I understand the circuits as a "Trotterization" of the temporal direction, instead of the spatial direction of the brick-wall case?
3-- In the introduction, the authors cited the quantum brickwork circuits [10] and [14-17]. Of course the authors knew very well, I find it useful to mention the poineer work in the 1980s [R4, R5] as a reference. I also think that the authors could cite one of my papers [R6], where a generalization of the brickwork construction (generalized to arbitrary periodicity) is considered. This comment is only suggestive.
4-- In section 2, when constructing Lax matrices, we obtained a (n x n) Lax matrix, where n is the dimension of the auxiliary space. I wonder if the dimension of the auxiliary space n is fixed by any parameter, such as the phase space dimension $n_X$ or $n_Y$.
5-- I wonder if the zero-curvature condition (2.2) is the usual zero-curvature condition rotated by $\pi/4$.
6-- The line above (2.10), $\Psi_{\tau} : X^L \times Y \to X^L \times Y$ seems to be contradictory with (2.8) and (2.9). For instance, Eq. (2.9) implies that $\Psi_{\tau} : Y \to Y$, and the right-hand side of (2.8) implies that $\Psi_{\tau} : X^{L-1} \times Y \to X^{L-1} \times Y$.
7-- I'm confused about the number of functionally independent charges in footnote 2. My understanding is that the total dimension of the phase space of $X^L$ is $L \cdot n_X$.
8-- In the set-theoretical Yang-Baxter equation part, the YBE written down is of difference form. I'm wondering if it is possible to generalize to the case that is not of difference form. (It seems to me that it should be possible.)
9-- On page 13 above Eq. (3.11), if the number of sites $L$ is finite, do we still need to take into account infinitely-many charges ${H_k}_{k=0}^{\infty}$? Can I understand Eq. (3.11) as a "classical generalized Gibbs ensemble"?
10-- It is not clear what the acronym "DST Lax matrix" stands for in footnote 6. It would be better if the authors use the full name or explain it.
11-- I'm not sure if Eqs. (4.16) and (4.17) do not contain typos. If I'm not mistaken, Eq. (4.16) should read $\alpha_{\ell-1} \nu_{\ell} = \mathcal{F}_{\ell-1} \nu_{\ell-1}$. Eq. (4.17) should read $\mathcal{F}_{\ell} = \begin{pmatrix} p_{\ell}+ \tau & - x_{\ell} / x_{\ell+1} \\ 1 & 0 \end{pmatrix} $. I urge the authors to double-check the formulae mentioned.
12-- I do not understand the statement below Eq. (4.19), namely why fixed point value $y^\ast$ must be positive and real. It would be great if the authors could explain.
13-- The continuous-time limit on page 16: Naïvely, I expect the continuous-time limit to be $\tau \to 0$, while the number of time steps goes to infinity. Would it arrive at the same result as the limit on page 16.
14-- On page 18 about the existence of fixed points, I do not understand the meaning of the index of fix points. Especially in the line under Eq. (4.52), the authors mentioned the possibility of having a fixed point with index two. What exactly is the index, and what is the physical meaning of having a fixed point of index two? The Lefschetz-Hopf theorem is purely topological. I'm wondering if the existence of fixed points is only determined by the topology of the target space of the classical sigma model.
15-- I'm wondering if the authors could comment on the physical meaning of the fixed points in terms of the many-body dynamics (any counterpart in the quantum case? any relation to the MPO fixed point?), and in the case of multiple fixed points, do we expect different physics when choosing different fixed points?
16-- It seems to me that if we choose a different choice of the random time-steps (e.g. non-Gaussian), we could obtain different "Brownian motions" than the Gaussian ones. I'm wondering if that is the correct understanding.
[R1] A.B. Zamolodchikov, "Fishing-net" diagrams as a completely integrable system, Phys. Lett. B, 97, 63-66 (1980).
[R2] Ö. Gürdoğan and V. Kazakov, New Integrable 4D Quantum Field Theories from Strongly Deformed Planar $\mathcal{N}=4$ Supersymmetric Yang-Mills Theory, Phys. Rev. Lett. 117, 201602 (2016).
[R3] D. Chicherin, V. Kazakov, F. Loebbert, D. Müller and D. Zhong, Yangian Symmetry for Fishnet Feynman Graphs, Phys. Rev. D 96, 121901 (2017).
[R4] C. Destri and H.J. De Vega, Light-cone lattice approach to fermionic theories in 2D: The massive Thirring model, Nucl. Phys. B 290, 363-391 (1987).
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Author: Žiga Krajnik on 2025-03-12 [id 5285]
(in reply to Report 1 by Yuan Miao on 2025-01-29)
We thank the referee for their detailed reading of the manuscript and many useful comments and suggestions which have improved the manuscript. Their comments are addressed below:
- Since there already exist another class of integrable circuits, widely known as brickwork (or brickwall) circuits, we have similarly sought a short and suggestive name for the proposed circuit architecture. We have decided to go with 'fishnet' circuits, compatibly with their diagrammatic representation. While we are aware of 'fishnet Feynmann diagrams' arising in supersymmetric gauge theories, we cannot recognize any immediate connections beyond the shared graph structure. Curiously, there likewise exist `brickwall' type diagrams. Given that the scope of our work bears no overlap with the studies of integrability in the domain of quantum field theories, we do not see much benefits of expanding on our choice of nomenclature.
- Contrary to the brickwork circuits which directly implement the Trotter-Suzuki decomposition of the evolution operator, the fishnet circuits cannot (at least directly) be interpreted in that way. Despite that, we show that fishnet circuits can be understood as an integrable time-space discretization of Hamiltonian (i.e. continuous-time) dynamics which, in sharp contrast to integrable brickwork circuits, exactly preserve the hierarchy of Hamiltonians for an arbitrary discrete time-step. To use the language of vertex models, the brickwork geometry corresponds to diagonal-to-diagonal transfer matrices whereas our fishnet construction employs row-to-row transfer matrices.
- We thank the referee for pointing out these references which have now been incorporated into the text.
- We do not think there is any general connection. For instance, it is known that Lax operators with different dimensions (of the auxiliary matrix space) can encode the same dynamics.
- Yes, the zero-curvature condition used in our construction is precisely the one used in construction of brickwork circuits rotated by $\pi/4$ (see Figure 1), apart from an extra permutation of arguments.
- Eqs. 2.8 and 2.9 involve a slight abuse of notation - they iteratively define a family of maps (distinguished by the number of their arguments) by passing the auxiliary variable through the chain, see Figure 2. The map $\Psi_\tau$ in Eq. 2.10 is obtained after $L$ iterations (i.e. after passing the auxiliary variable across the entire length of the chain). We afforded this slight imprecision to avoid using unnecessarily heavy notation. To avoid confusion we have now added an additional subscript to distinguish these maps in 2.8 and 2.9, and similarly also for 2.14 and 2.15.
- We thank the referee for catching this typo which has been corrected.
- Yes, we believe a similar construction can be accomplished using a non-difference form of the Yang-Baxter relation provided the analogous non-difference zero-curvature condition can be resolved explicitly. That said, some of the outlined properties in the current construction may no longer hold.
- Yes, Eq. 3.11 can be indeed be viewed a classical generalized Gibbs ensemble. Notice however that the question of completeness of classical generalized Gibbs ensembles is not a trivial one, and it presently remains unclear clear how to identify and construct a complete set of charges that characterize local maximum-entropy ensembles.
- The acronym DST stands for `discrete self-trapping'. Its single occurrence has now been eliminated.
- We thank the referee for catching the mistake in 4.17 which has now been corrected. We have also made several other minor modifications in the paragraph regarding the fixed points of the Toda lattice to improve readability.
- Note that Eq. 4.8 gives $x^\prime_{\ell} = y_{\ell-1}$. Specializing to $\ell=1$ immediately gives that the fixed point $y^*$ must itself lie in the the phase space of the Toda lattice.
- One needs to take $\tau \to \infty$ to obtain the continuous-time limit, taking $\tau \to 0$ does not recover the correct limit. Note that this is one particularily of the Toda lattice; to obtain the local conserved quantities one likewise has to expand the transfer matrix around $\lambda \to \infty$.
- The index associated to a fixed point of a function counts the degeneracy of the fixed point under a generic perturbation of that function, see e.g. Chapter 3 of Differential Topology by V. Guillemin and A. Pollack. In the updated version we have corrected a mistake in the discussion of the number of fixed points. In particular, the index of a fixed point can be negative so that the Lefschtz-Hopf theorem only gives a lower bound on the number of fixed points. More concretely, the index of a fixed point is related to the topological behavior of the function in the vicinity of the fixed point. In two dimensions non-degenerate attractive and repulsive fixed points have index $1$ while non-degenerate saddle fixed points have index $-1$. However, since these considerations are not very essential to the manuscript (in fact, we also obtain the fixed points analytically), we now only note that the existence of a fixed point is guaranteed without discussing their multiplicity. In addition to topological properties of the phase-space, the existence of fixed points also depends on the homology properties of the local two-body map. In the present case of the 2-sphere, the relevant property is orientation preservation (or lack thereof) which determines the trace of the map induced on the second homology group. As noted, the map we consider is orientation-preserving whence the existence of a fixed point follows by the Lefschetz-Hopf theorem. On the other hand, an orientation-reversing map will produce a vanishing Lefschetz number and the existence of a fixed point is not guaranteed, as is easily seen by considering sphere inversion, i.e. the map $x \mapsto -x$ on the 2-sphere.
- The quantum analogue of the fixed-point condition would be to trace out the auxiliary (i.e. horizontal) space. However, it is presently not evident how the classical fixed-point condition arises in the semi-classical limit of the corresponding integrable quantum model (as the auxiliary dimension grows large). While different fixed points define distinct dynamics we have not encountered any meaningful physical differences, at least not in the considered cases. This is perhaps not surprising since the resulting dynamics always preserved the transfer matrix of the model. On the other hand, we cannot exclude that models with a richer phase-space and fixed-point structure might exhibit different behavior. We believe these questions merits further attention.
- The referee's intuition is correct. Our results were obtained assuming a finite variance of the time-step distribution. Relaxing this assumption and varying the tail exponent of the time-step distribution results in continuous-time dynamics with a smoothly varying dynamical exponent generated by Levy stable distributions by the generalized central limit theorem.
Author: Žiga Krajnik on 2025-03-12 [id 5288]
(in reply to Report 3 on 2025-02-06)We thank the referee for their reading and evaluation of our work. As noted by the referee, the isotropic two-body spin map is already known. The core idea of this manuscript is, however, to outline how maps of this type can be used to construct integrable circuits with a number of distinct properties. The isotropic spin propagator is merely a widely known and convenient example used to illustrate the general scheme. Their comments are addressed below:
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