Integrable fishnet circuits and Brownian solitons

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:

1. In our understanding, the constructed discrete-time dynamics admits an interpolating Hamiltonian flow generated by some linear combination of commuting Hamiltonians of the continuum theory, as can be readily deduced by the preservation of the classical spectral curve. This explains our introduction of commuting flows. Consequently, the dynamics remain confined to invariant tori of the Hamiltonian counterpart, and the elementary one-step propagator acts simply by discrete shifts (depending on $\tau$) of canonical angle variables. We nonetheless decided to not delve into these aspects (which would require us to introduce the framework of separation of variables) in order to not deviate too much from the main scope of the paper.
2. The Poisson structure of physical and auxiliary degrees of freedom is indeed prescribed by the $r$-matrix in Eq. (2.1). Concerning canonicity, we believe that there should exist a fairly general (i.e. model-independent) way of establishing this property, but it has so far eluded us. We thus proceed with a case-by-base analysis. The main difference between the examples of Toda and Landau-Lifschitz is that the auxiliary and physical Lax operators are inequivalent in the former case despite sharing the same $r$-matrix. There is no a priori guarantee that $\psi_\tau$, the solution to the local zero-curvature condition Eq. (2.2), represents a canonical map. Indeed, the local Toda propagator Eqs. (4.8-4.11) is an example of a non-cannonical solution. Evidently it then follows that the composite propagators $\Psi^\pm_\tau$ are also not canonical. The physical Toda propagator $\Phi^\pm_\tau$ however becomes canonical after eliminating the auxiliary variable by periodically closing the chain. The solution of the local discrete zero-curvature condition for the Toda lattice is well-known, and to our knowledge was first obtained by Sklyanin who also obtained the equivalent of $\Phi^-_\tau$ and demonstrated its canonicity, as noted in the manuscript, see Refs. [23] and [24]. Our construction shows that the same type of dynamics can be realized for other integrable models with a solution of the discrete zero-curvature condition given as the input. We have also established a number of nontrivial properties of such dynamics in a model-independent manner. Among others, we demonstrate how by alternating left and right propagators (in absence of inhomogeneities) one can remedy the undesired chirality of the elementary one-step propagators. The auxiliary and physical Lax operators of the Landau-Lifschitz are isomorphic. The solutions $\psi_\tau$ of the corresponding discrete zero-curvature conditions are known (see Refs. [11-13]) where it is also demonstrated that they preserve the Poisson structure of Eq. (2.1), from which it immediately shows that they are canonical maps. Note that explicit expressions for Lax operators or propagators in canonical coordinates are unnecessary to obtain this result. Since $\psi_\tau$ is canonical this directly gives that $\Psi_\tau$ is also canonical as we note after Eq. (4.47). It then remains to understand whether eliminating the auxiliary (i.e. horizontal) variable by imposing periodicity spoils the canonicity of the resulting map $\Phi^-_\tau$. The last question is addressed in Appendix A in a model-independent manner. In particular, we show that starting with a canonical map $\Psi_\tau$ and eliminating the auxiliary variables by imposing the periodicty condition Eq. (2.10) necessarily yields a canonical map $\Phi_\tau^-$. Given that our example of the Landau-Lifschitz model produces a canonical $\Psi_\tau$ the proof of Appendix A applies, while the Toda example gives a non-canonical $\Psi_\tau$, hence the reference to Sklyanin's proof.
3. It appears that there are certain misunderstandings that require some clarification. Firstly, stochasticity enters our construction via the random choice of time-step parameters and thus concerns the time-evolution and not initial conditions. This complies with the conventional use of `stochastic dynamics' in the realm of many-body systems, where randomness enter through the equation of motion. Initial condition are of course random, since we deal with maximal entropy ensembles. But we could just as easily have studied a quench protocol by prescribing some fixed initial configuration. Secondly, the propagator given by equation (5.5) is integrable! Its integrability is guaranteed by the construction outlined in Sections 2 and 3. Crucially, randomness in the time-step parameters does not break integrability. This indeed is the key distinction to integrable brickwork circuits, where picking different $\tau$ along the evolution invariably spoils integrability. In fishnet circuits, exact conservation of the charges, stated in Eq. (3.7), follows from the conservation of the transfer functions. Consequently, if by 'energy' one has in mind one of the conserved quantities as per Eq. (3.7), there are no dynamics.

Minor commnets:

1. The symbol $\simeq$ indicates proportionality up to a scalar factor as we now state explicitly. As correctly observed by the referee, the Toda Lax operators do not satisfy the inversion relation. Note, however, that the inverse relation is not required for the construction. We made a different statement, namely if Eq. (2.39) is satisfied, then the inverse propagator can be realized by negating the time-step, see Eq. (2.40).
2. To realize the left propagator in Figure 3 in terms of the right propagator with a reversed direction and a negative time-step we use the inversion relation (2.39) as noted in Eq. (2.42) and the caption, hence its inclusion in the caption of Figure 3.
3. It is well known that the local conserved quantities of the Toda lattice are obtained by expanding the transfer matrix around $\lambda \to \infty$ (in the standard parametrization of the spectral plane). To retrieve the corresponding continuous-time dynamics it is necessary to let $\tau \to \infty$. As we explain after Eq. (4.33), the doubling of the propagator is necessary to bypass the staggering of signs which obstructs a smooth continuous limit.
4. We have not expressed ourselves clearly. We are well aware of previous works that have succeeded in deriving the exact transport coefficients (Onsager matrix) governing diffusion-scale dynamics in integrable models. We implicitly had in mind something else: the mechanism for diffusive (Brownian) dynamics outlined in our work is distinct from that treated in the aforementioned works, and presently it is not obvious how to incorporate it within the existing framework of generalized hydrodynamics. The origin of diffusive-scale dynamics in our fishnet circuits is not `diffusion by convention' as in deterministic interacting integrable systems but rather stochasticity in time which we describe by introducing stochastic dressing.

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:

1. We thank the referee for point out this typo which has been corrected.
2. 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.
3. 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.
4. We now explicitly mention that the transfer matrices $\mathcal{T}^{(n)}$ are obtained by tracing powers of the monodromy matrix before Eq. 1.5.
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.
6. 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.
7. We thank the referee for catching this typo which has been corrected.
8. We thank the referee for point this out. We have corrected Eq. 2.21 so that it also holds for $n,m \geq 1$.
9. 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 $\lambda \to \infty$.
10. 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.
11. 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].
12. 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.
13. We now explicitly define $\sigma^0$ after Eq. 4.7 to avoid confusion.
14. 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.
15. 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.
16. We now use a different font to denote the homology groups.
17. 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.
18. 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.

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:

1. We thank the referee for point out this typo which has been corrected.
2. 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.
3. 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.
4. We now explicitly mention that the transfer matrices $\mathcal{T}^{(n)}$ are obtained by tracing powers of the monodromy matrix before Eq. 1.5.
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.
6. 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.
7. We thank the referee for catching this typo which has been corrected.
8. We thank the referee for point this out. We have corrected Eq. 2.21 so that it also holds for $n,m \geq 1$.
9. 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 $\lambda \to \infty$.
10. 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.
11. 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].
12. 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.
13. We now explicitly define $\sigma^0$ after Eq. 4.7 to avoid confusion.
14. 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.
15. 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.
16. We now use a different font to denote the homology groups.
17. 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.
18. 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.

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:

1. 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.
2. 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.
3. We thank the referee for pointing out these references which have now been incorporated into the text.
4. 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.
5. 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.
6. 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.
7. We thank the referee for catching this typo which has been corrected.
8. 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.
9. 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.
10. The acronym DST stands for `discrete self-trapping'. Its single occurrence has now been eliminated.
11. 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.
12. 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.
13. 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 $\lambda \to \infty$.
14. 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.
15. 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.
16. 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.