The Stochastic Resonance Program (Part 1)

10 May, 2014

guest post by David Tanzer

At the Azimuth Code Project, we are aiming to produce educational software that is relevant to the Earth sciences and the study of climate. Our present software takes the form of interactive web pages, which allow you to experiment with the parameters of models and view their outputs. But to fully understand the meaning of a program, we need to know about the concepts and theories that inform it. So we will be writing articles to explain both the programs themselves and the math and science behind them.

In this two-part series, I’ll explain this program:

Stochastic resonance.

Check it out—it runs on your browser! It was created by Allan Erskine and Glyn Adgie. In the Azimuth blog article Increasing the Signal-to-Noise Ratio with More Noise, Glyn Adgie and Tim van Beek give a nice explanation of the idea of stochastic resonance, which includes some clear and exciting graphs.

My goal today is give a compact, developer-oriented introduction to stochastic resonance, which will set the context for the next blog article, where I’ll dissect the program itself. By way of introduction, I am a software developer with research training in computer science. It’s a new area for me, and any clarifications will be welcome!

The concept of stochastic resonance

Stochastic resonance is a phenomenon, occurring under certain circumstances, in which a noise source may amplify the effect of a weak signal. This concept was used in an early hypothesis about the timing of ice-age cycles, and has since been applied to a wide range of phenomena, including neuronal detection mechanisms and patterns of traffic congestion.

Suppose we have a signal detector whose internal, analog state is driven by an input signal, and suppose the analog states are partitioned into two regions, called “on” and “off” — this is a digital state, abstracted from the analog state. With a light switch, we could take the force as the input signal, the angle as the analog state, and the up/down classification of the angle as the digital state.

Consider the effect of a periodic input signal on the digital state. Suppose the wave amplitude is not large enough to change the digital state, yet large enough to drive the analog state close to the digital state boundary. Then, a bit of random noise, occurring near the peak of an input cycle, may “tap” the system over to the other digital state. So we will see a probability of state-transitions that is synchronized with the input signal. In a complex way, the noise has amplified the input signal.

But it’s a pretty funky amplifier! Here is a picture from the Azimuth library article on stochastic resonance:

Stochastic resonance has been found in the signal detection mechanisms of neurons. There are, for example, cells in the tails of crayfish that are tuned to low-frequency signals in the water caused by predator motions. These signals are too weak to cross the firing threshold for the neurons, but with the right amount of noise, they do trigger the neurons.

See:

Stochastic resonance, Azimuth Library.

Stochastic resonance in neurobiology, David Lyttle.

Bistable stochastic resonance and Milankovitch theories of ice-age cycles

Stochastic resonance was originally formulated in terms of systems that are bistable — where each digital state is the basin of attraction of a stable equilibrium.

An early application of stochastic resonance was to a hypothesis, within the framework of bistable climate dynamics, about the timing of the ice-age cycles. Although it has not been confirmed, it remains of interest (1) historically, (2) because the timing of ice-age cycles remains an open problem, and (3) because the Milankovitch hypothesis upon which it rests is an active part of the current research.

In the bistable model, the climate states are a cold, “snowball” Earth and a hot, iceless Earth. The snowball Earth is stable because it is white, and hence reflects solar energy, which keeps it frozen. The iceless Earth is stable because it is dark, and hence absorbs solar energy, which keeps it melted.

The Milankovitch hypothesis states that the drivers of climate state change are long-duration cycles in the insolation — the solar energy received in the northern latitudes — caused by periodic changes in the Earth’s orbital parameters. The north is significant because that is where the glaciers are concentrated, and so a sufficient “pulse” in northern temperatures could initiate a state change.

Three relevant astronomical cycles have been identified:

• Changing of the eccentricity of the Earth’s elliptical orbit, with a period of 100 kiloyears

• Changing of the obliquity (tilt) of the Earth’s axis, with a period of 41 kiloyears

• Precession (swiveling) of the Earth’s axis, with a period of 23 kiloyears

In the stochastic resonance hypothesis, the Milankovitch signal is amplified by random events to produce climate state changes. In more recent Milankovitch theories, a deterministic forcing mechanism is used. In a theory by Didier Paillard, the climate is modeled with three states, called interglacial, mild glacial and full glacial, and the state changes depend on the volume of ice as well as the insolation.

See:

Milankovitch cycle, Azimuth Library.

Mathematics of the environment (part 10), John Baez. This gives an exposition of Paillard’s theory.

Bistable systems defined by a potential function

Any smooth function with two local minima can be used to define a bistable system. For instance, consider the function V(x) = x^4/4 - x^2/2:

To define the bistable system, construct a differential equation where the time derivative of x is set to the negative of the derivative of the potential at x:

dx/dt = -V'(x) = -x^3 + x = x(1 - x^2)

So, for instance, where the potential graph is sloping upward as x increases, -V'(x) is negative, and this sends X(t) ‘downhill’ towards the minimum.

The roots of V'(x) yield stable equilibria at 1 and -1, and an unstable equilibrium at 0. The latter separates the basins of attraction for the stable equilibria.

Discrete stochastic resonance

Now let’s look at a discrete-time model which exhibits stochastic resonance. This is the model used in the Azimuth demo program.

We construct the discrete-time derivative, using the potential function, a sampled sine wave, and a normally distributed random number:

\Delta X_t = -V'(X_t) * \Delta t + \mathrm{Wave}(t) + \mathrm{Noise}(t) =
X_t (1 - X_t^2) \Delta t + \alpha * \sin(\omega t) + \beta * \mathrm{GaussianSample}(t)

where \Delta t is a constant and t is restricted to multiples of \Delta t.

This equation is the discrete-time counterpart to a continuous-time stochastic differential equation.

Next time, we will look into the Azimuth demo program itself.


Quantum Frontiers in Network Science

6 May, 2014

guest post by Jacob Biamonte

There’s going to be a workshop on quantum network theory in Berkeley this June. The event is being organized by some of my collaborators and will be a satellite of the biggest annual network science conference, NetSci.

A theme of the Network Theory series here on Azimuth has been to merge ideas appearing in quantum theory with other disciplines. Remember the first post by John which outlined the goal of a general theory of networks? Well, everyone’s been chipping away at this stuff for a few years now and I think you’ll agree that this workshop seems like an excellent way to push these topics even further, particularly as they apply to complex networks.

The event is being organized by Mauro Faccin, Filippo Radicchi and Zoltán Zimborás. You might recall when Tomi Johnson first explained to us some ideas connecting quantum physics with the concepts of complex networks (see Quantum Network Theory Part 1 and Part 2). Tomi’s going to be speaking at this event. I understand there is even still a little bit of space left to contribute talks and/or to attend. I suspect that those interested can sort this out by emailing the organizers or just follow the instructions to submit an abstract.

They have named their event Quantum Frontiers in Network Science or QNET for short. Here’s their call.

Quantum Frontiers in Network Science

This year the biggest annual network science conference, NetSci will take place in Berkeley California on 2-6 June. We are organizing a one-day Satellite Workshop on Quantum Frontiers in Network Science (QNET).

quantum netsci2014

A grand challenge in contemporary complex network science is to reconcile the staple “statistical mechanics based approach” with a theory based on quantum physics. When considering networks where quantum coherence effects play a non-trivial role, the predictive power of complex network science has been shown to break down. A new theory is now being developed which is based on quantum theory, from first principles. Network theory is a diverse subject which developed independently in several disciplines to rely on graphs with additional structure to model complex systems. Network science has of course played a significant role in quantum theory, for example in topics such as tensor network states, chiral quantum walks on complex networks, categorical tensor networks, and categorical models of quantum circuits, to name only a few. However, the ideas of complex network science are only now starting to be united with modern quantum theory. From this respect, one aim of the workshop is to put in contact two big and generally not very well connected scientific communities: statistical and quantum physicists.

The topic of network science underwent a revolution when it was realized that systems such as social or transport networks could be interrelated through common network properties, but what are the relevant properties to consider when facing quantum systems? This question is particularly timely as there has been a recent push towards studying increasingly larger quantum mechanical systems, where the analysis is only beginning to undergo a shift towards embracing the concepts of complex networks.

brain network

For example, theoretical and experimental attention has turned to explaining transport in photosynthetic complexes comprising tens to hundreds of molecules and thousands of atoms using quantum mechanics. Likewise, in condensed matter physics using the language of “chiral quantum walks”, the topological structure of the interconnections comprising complex materials strongly affects their transport properties.

An ultimate goal is a mathematical theory and formal description which pinpoints the similarities and differences between the use of networks throughout the quantum sciences. This would give rise to a theory of networks augmenting the current statistical mechanics approach to complex network structure, evolution, and process with a new theory based on quantum mechanics.

Topics of special interest to the satellite include

• Quantum transport and chiral quantum walks on complex networks
• Detecting community structure in quantum systems
• Tensor algebra and multiplex networks
• Quantum information measures (such as entropy) applied to complex networks
• Quantum critical phenomena in complex networks
• Quantum models of network growth
• Quantum techniques for reaction networks
• Quantum algorithms for problems in complex network science
• Foundations of quantum theory in relation to complex networks and processes thereon
• Quantum inspired mathematics as a foundation for network science

Info

QNET will be held at the NetSci Conference venue at the Clark Kerr Campus of the University of California, on June 2nd in the morning (8am-1pm).

Links

• Main conference page: NetSci2014
Call for abstracts and the program

It sounds interesting! You’ll notice that the list of topics seems reminiscent of some of the things we’ve been talking about right here on Azimuth! A general theme of the Network Theory Series has been geared towards developing frameworks to describe networked systems through a common language and then to map the use of tools and results across disciplines. It seems like a great place to talk about these ideas. Oh, and here’s a current list of the speakers:

Leonardo Banchi (UCL, London)
Ginestra Bianconi (London)
Silvano Garnerone (IQC, Waterloo)
Laetitia Gauvin (ISI Foundation)
Marco Javarone (Sassari)
Tomi Johnson (Oxford)

and again, the organizers are

Mauro Faccin (ISI Foundation)
Filippo Radicchi (Indiana University)
Zoltán Zimborás (UCL)

From the call, we can notice that a central discussion topic at QNET will be about contrasting stochastic and quantum mechanics. Here on Azimuth we like this stuff. You might remember that stochastic mechanics was formulated in the network theory series to mathematically resemble quantum theory (see e.g. Part 12). This formalism was then employed to produce several results, including a stochastic version of Noether’s theorem by John and Brendan in Parts 11 and 13—recently Ville has also written Noether’s Theorem: Quantum vs Stochastic. Several other results were produced by relating quantum field theory to Petri nets from population biology and to chemical reaction networks in chemistry (see the Network Theory homepage). It seems to me that people attending QNET will be interested in these sorts of things, as well as other related topics.

One of the features of complex network science is that it is often numerically based and geared directly towards interesting real-world applications. I suspect some interesting results should stem from the discussions that will take place at this workshop.

By the way, here’s a view of downtown San Francisco at dusk from Berkeley Hills California from the NetSci homepage:

San Francisco

Noether’s Theorem: Quantum vs Stochastic

3 May, 2014

guest post by Ville Bergholm

In 1915 Emmy Noether discovered an important connection between the symmetries of a system and its conserved quantities. Her result has become a staple of modern physics and is known as Noether’s theorem.

Photo of Emmy Noether

The theorem and its generalizations have found particularly wide use in quantum theory. Those of you following the Network Theory series here on Azimuth might recall Part 11 where John Baez and Brendan Fong proved a version of Noether’s theorem for stochastic systems. Their result is now published here:

• John Baez and Brendan Fong, A Noether theorem for stochastic mechanics, J. Math. Phys. 54:013301 (2013).

One goal of the network theory series here on Azimuth has been to merge ideas appearing in quantum theory with other disciplines. John and Brendan proved their stochastic version of Noether’s theorem by exploiting ‘stochastic mechanics’ which was formulated in the network theory series to mathematically resemble quantum theory. Their result, which we will outline below, was different than what would be expected in quantum theory, so it is interesting to try to figure out why.

Recently Jacob Biamonte, Mauro Faccin and myself have been working to try to get to the bottom of these differences. What we’ve done is prove a version of Noether’s theorem for Dirichlet operators. As you may recall from Parts 16 and 20 of the network theory series, these are the operators that generate both stochastic and quantum processes. In the language of the series, they lie in the intersection of stochastic and quantum mechanics. So, they are a subclass of the infinitesimal stochastic operators considered in John and Brendan’s work.

The extra structure of Dirichlet operators—compared with the wider class of infinitesimal stochastic operators—provided a handle for us to dig a little deeper into understanding the intersection of these two theories. By the end of this article, astute readers will be able to prove that Dirichlet operators generate doubly stochastic processes.

Before we get into the details of our proof, let’s recall first how conservation laws work in quantum mechanics, and then contrast this with what John and Brendan discovered for stochastic systems. (For a more detailed comparison between the stochastic and quantum versions of the theorem, see Part 13 of the network theory series.)

The quantum case

I’ll assume you’re familiar with quantum theory, but let’s start with a few reminders.

In standard quantum theory, when we have a closed system with n states, the unitary time evolution of a state |\psi(t)\rangle is generated by a self-adjoint n \times n matrix H called the Hamiltonian. In other words, |\psi(t)\rangle satisfies Schrödinger’s equation:

i \hbar \displaystyle{\frac{d}{d t}} |\psi(t) \rangle = H |\psi(t) \rangle.

The state of a system starting off at time zero in the state |\psi_0 \rangle and evolving for a time t is then given by

|\psi(t) \rangle = e^{-i t H}|\psi_0 \rangle.

The observable properties of a quantum system are associated with self-adjoint operators. In the state |\psi \rangle, the expected value of the observable associated to a self-adjoint operator O is

\langle O \rangle_{\psi} = \langle \psi | O | \psi \rangle

This expected value is constant in time for all states if and only if O commutes with the Hamiltonian H:

[O, H] = 0 \quad \iff \quad \displaystyle{\frac{d}{d t}} \langle O \rangle_{\psi(t)} = 0 \quad \forall \: |\psi_0 \rangle, \forall t.

In this case we say O is a ‘conserved quantity’. The fact that we have two equivalent conditions for this is a quantum version of Noether’s theorem!

The stochastic case

In stochastic mechanics, the story changes a bit. Now a state |\psi(t)\rangle is a probability distribution: a vector with n nonnegative components that sum to 1. Schrödinger’s equation gets replaced by the master equation:

\displaystyle{\frac{d}{d t}} |\psi(t) \rangle = H |\psi(t) \rangle

If we start with a probability distribution |\psi_0 \rangle at time zero and evolve it according to this equation, at any later time have

|\psi(t)\rangle = e^{t H} |\psi_0 \rangle.

We want this always be a probability distribution. To ensure that this is so, the Hamiltonian H must be infinitesimal stochastic: that is, a real-valued n \times n matrix where the off-diagonal entries are nonnegative and the entries of each column sum to zero. It no longer needs to be self-adjoint!

When H is infinitesimal stochastic, the operators e^{t H} map the set of probability distributions to itself whenever t \ge 0, and we call this family of operators a continuous-time Markov process, or more precisely a Markov semigroup.

In stochastic mechanics, we say an observable O is a real diagonal n \times n matrix, and its expected value is given by

\langle O\rangle_{\psi} = \langle \hat{O} | \psi \rangle

where \hat{O} is the vector built from the diagonal entries of O. More concretely,

\langle O\rangle_{\psi} = \displaystyle{ \sum_i O_{i i} \psi_i }

where \psi_i is the ith component of the vector |\psi\rangle.

Here is a version of Noether’s theorem for stochastic mechanics:

Noether’s Theorem for Markov Processes (Baez–Fong). Suppose H is an infinitesimal stochastic operator and O is an observable. Then

[O,H] =0

if and only if

\displaystyle{\frac{d}{d t}} \langle O \rangle_{\psi(t)} = 0

and

\displaystyle{\frac{d}{d t}}\langle O^2 \rangle_{\psi(t)} = 0

for all t \ge 0 and all \psi(t) obeying the master equation.   █

So, just as in quantum mechanics, whenever [O,H]=0 the expected value of O will be conserved:

\displaystyle{\frac{d}{d t}} \langle O\rangle_{\psi(t)} = 0

for any \psi_0 and all t \ge 0. However, John and Brendan saw that—unlike in quantum mechanics—you need more than just the expectation value of the observable O to be constant to obtain the equation [O,H]=0. You really need both

\displaystyle{\frac{d}{d t}} \langle O\rangle_{\psi(t)} = 0

together with

\displaystyle{\frac{d}{d t}} \langle O^2\rangle_{\psi(t)} = 0

for all initial data \psi_0 to be sure that [O,H]=0.

So it’s a bit subtle, but symmetries and conserved quantities have a rather different relationship than they do in quantum theory.

A Noether theorem for Dirichlet operators

But what if the infinitesimal generator of our Markov semigroup is also self-adjoint? In other words, what if H is both an infinitesimal stochastic matrix but also its own transpose: H = H^\top? Then it’s called a Dirichlet operator… and we found that in this case, we get a stochastic version of Noether’s theorem that more closely resembles the usual quantum one:

Noether’s Theorem for Dirichlet Operators. If H is a Dirichlet operator and O is an observable, then

[O, H] = 0 \quad \iff \quad \displaystyle{\frac{d}{d t}} \langle O \rangle_{\psi(t)} = 0 \quad \forall \: |\psi_0 \rangle, \forall t \ge 0

Proof. The \Rightarrow direction is easy to show, and it follows from John and Brendan’s theorem. The point is to show the \Leftarrow direction. Since H is self-adjoint, we may use a spectral decomposition:

H = \displaystyle{ \sum_k E_k |\phi_k \rangle \langle \phi_k |}

where \phi_k are an orthonormal basis of eigenvectors, and E_k are the corresponding eigenvalues. We then have:

\displaystyle{\frac{d}{d t}} \langle O \rangle_{\psi(t)} = \langle \hat{O} | H e^{t H} |\psi_0 \rangle = 0 \quad \forall \: |\psi_0 \rangle, \forall t \ge 0

\iff \quad \langle \hat{O}| H e^{t H} = 0 \quad \forall t \ge 0

\iff \quad \sum_k \langle \hat{O} | \phi_k \rangle E_k e^{t E_k} \langle \phi_k| = 0 \quad \forall t \ge 0

\iff \quad \langle \hat{O} | \phi_k \rangle E_k e^{t E_k} = 0 \quad \forall t \ge 0

\iff \quad |\hat{O} \rangle \in \mathrm{Span}\{|\phi_k \rangle \, : \; E_k = 0\} = \ker \: H,

where the third equivalence is due to the vectors |\phi_k \rangle being linearly independent. For any infinitesimal stochastic operator H the corresponding transition graph consists of m connected components iff we can reorder (permute) the states of the system such that H becomes block-diagonal with m blocks. Now it is easy to see that the kernel of H is spanned by m eigenvectors, one for each block. Since H is also symmetric, the elements of each such vector can be chosen to be ones within the block and zeros outside it. Consequently

|\hat{O} \rangle \in \ker \: H

implies that we can choose the basis of eigenvectors of O to be the vectors |\phi_k \rangle, which implies

[O, H] = 0

Alternatively,

|\hat{O} \rangle \in \ker \, H

implies that

|\hat{O^2} \rangle \in \ker \: H \; \iff \; \cdots \; \iff \; \displaystyle{\frac{d}{d t}} \langle O^2 \rangle_{\psi(t)} = 0 \; \forall \: |\psi_0 \rangle, \forall t \ge 0,

where we have used the above sequence of equivalences backwards. Now, using John and Brendan’s original proof, we can obtain [O, H] = 0.   █

In summary, by restricting ourselves to the intersection of quantum and stochastic generators, we have found a version of Noether’s theorem for stochastic mechanics that looks formally just like the quantum version! However, this simplification comes at a cost. We find that the only observables O whose expected value remains constant with time are those of the very restricted type described above, where the observable has the same value in every state in a connected component.

Puzzles

Suppose we have a graph whose graph Laplacian matrix H generates a Markov semigroup as follows:

U(t) = e^{t H}

Puzzle 1. Suppose that also H = H^\top, so that H is a Dirichlet operator and hence i H generates a 1-parameter unitary group. Show that the indegree and outdegree of any node of our graph must be equal. Graphs with this property are called balanced.

Puzzle 2. Suppose that U(t) = e^{t H} is doubly stochastic Markov semigroup, meaning that for all t \ge 0 each row and each column of U(t) sums to 1:

\displaystyle{ \sum_i U(t)_{i j} = \sum_j U(t)_{i j} = 1 }

and all the matrix entries are nonnegative. Show that the Hamiltonian H obeys

\displaystyle{\sum_i H_{i j} = \sum_j H_{i j} = 0 }

and all the off-diagonal entries of H are nonnegative. Show the converse is also true.

Puzzle 3. Prove that any doubly stochastic Markov semigroup U(t) is of the form e^{t H} where H is the graph Laplacian of a balanced graph.

Puzzle 4. Let O(t) be a possibly time-dependent observable, and write \langle O(t) \rangle_{\psi(t)} for its expected value with respect to some initial state \psi_0 evolving according to the master equation. Show that

\displaystyle{ \frac{d}{d t}\langle O(t)\rangle_{\psi(t)} = \left\langle [O(t), H] \right\rangle_{\psi(t)} + \left\langle \frac{\partial O(t)}{\partial t}\right\rangle_{\psi(t)} }

This is a stochastic version of the Ehrenfest theorem.


Networks of Dynamical Systems

18 March, 2014

guest post by Eugene Lerman

Hi, I’m Eugene Lerman. I met John back in the mid 1980s when John and I were grad students at MIT. John was doing mathematical physics and I was studying symplectic geometry. We never talked about networks. Now I teach in the math department at the University of Illinois at Urbana, and we occasionally talk about networks on his blog.

A few years ago a friend of mine who studies locomotion in humans and other primates asked me if I knew of any math that could be useful to him.

I remember coming across an expository paper on ‘coupled cell networks’:

• Martin Golubitsky and Ian Stewart, Nonlinear dynamics of networks: the groupoid formalism, Bull. Amer. Math. Soc. 43 (2006), 305–364.

In this paper, Golubitsky and Stewart used the study of animal gaits and models for the hypothetical neural networks called ‘central pattern generators’ that give rise to these gaits to motivate the study of networks of ordinary differential equations with symmetry. In particular they were interested in ‘synchrony’. When a horse trots, or canters, or gallops, its limbs move in an appropriate pattern, with different pairs of legs moving in synchrony:


They explained that synchrony (and the patterns) could arise when the differential equations have finite group symmetries. They also proposed several systems of symmetric ordinary differential equations that could generate the appropriate patterns.

Later on Golubitsky and Stewart noticed that there are systems of ODEs that have no group symmetries but whose solutions nonetheless exhibit certain synchrony. They found an explanation: these ODEs were ‘groupoid invariant’. I thought that it would be fun to understand what ‘groupoid invariant’ meant and why such invariance leads to synchrony.

I talked my colleague Lee DeVille into joining me on this adventure. At the time Lee had just arrived at Urbana after a postdoc at NYU. After a few years of thinking about these networks Lee and I realized that strictly speaking one doesn’t really need groupoids for these synchrony results and it’s better to think of the social life of networks instead. Here is what we figured out—a full and much too precise story is here:

• Eugene Lerman and Lee DeVille, Dynamics on networks of manifolds.

Let’s start with an example of a class of ODEs with a mysterious property:

Example. Consider this ordinary differential equation for a function \vec{x} : \mathbb{R} \to {\mathbb{R}}^3

\begin{array}{rcl}  \dot{x}_1&=& f(x_1,x_2)\\  \dot{x}_2&=& f(x_2,x_1)\\  \dot{x}_3&=& f(x_3, x_2)  \end{array}

for some function f:{\mathbb{R}}^2 \to {\mathbb{R}}. It is easy to see that a function x(t) solving

\displaystyle{  \dot{x} = f(x,x)  }

gives a solution of these equations if we set

\vec{x}(t) = (x(t),x(t),x(t))

You can think of the differential equations in this example as describing the dynamics of a complex system built out of three interacting subsystems. Then any solution of the form

\vec{x}(t) = (x(t),x(t),x(t))

may be thought of as a synchronization: the three subsystems are evolving in lockstep.

One can also view the result geometrically: the diagonal

\displaystyle{  \Delta = \{(x_1,x_2, x_3)\in {\mathbb{R}}^3 \mid x_1 =x_2 = x_3\}  }

is an invariant subsystem of the continuous-time dynamical system defined by the differential equations. Remarkably enough, such a subsystem exists for any choice of a function f.

Where does such a synchronization or invariant subsystem come from? There is no apparent symmetry of {\mathbb{R}}^3 that preserves the differential equations and fixes the diagonal \Delta, and thus could account for this invariant subsystem. It turns out that what matters is the structure of the mutual dependencies of the three subsystems making up the big system. That is, the evolution of x_1 depends only on x_1 and x_2, the evolution of x_2 depends only on x_2 and x_3, and the evolution of x_3 depends only on x_3 and x_2.

These dependencies can be conveniently pictured as a directed graph:

The graph G has no symmetries. Nonetheless, the existence of the invariant subsystem living on the diagonal \Delta can be deduced from certain properties of the graph G. The key is the existence of a surjective map of graphs

\varphi :G\to G'

from our graph G to a graph G' with exactly one node, call it a, and one arrow. It is also crucial that \varphi has the following lifting property: there is a unique way to lift the one and only arrow of G' to an arrow of G once we specify the target node of the lift.

We now formally define the notion of a ‘network of phase spaces’ and of a continuous-time dynamical system on such a network. Equivalently, we define a ‘network of continuous-time dynamical systems’. We start with a directed graph

G=\{G_1\rightrightarrows G_0\}

Here G_1 is the set of edges, G_0 is the set of nodes, and the two arrows assign to an edge its source and target, respectively. To each node we attach a phase space (or more formally a manifold, perhaps with boundary or corners). Here ‘attach’ means that we choose a function {\mathcal P}:G_0 \to {\mathsf{PhaseSpace}}; it assigns to each node a\in G_0 a phase space {\mathcal P}(a).

In our running example, to each node of the graph G we attach the real line {\mathbb{R}}. (If we think of the set G_0 as a discrete category and {\mathsf{PhaseSpace}} as a category of manifolds and smooth maps, then {\mathcal P} is simply a functor.)

Thus a network of phase spaces is a pair (G,{\mathcal P}), where G is a directed graph and {\mathcal P} is an assignment of phase spaces to the nodes of G.

We think of the collection \{{\mathcal P}(a)\}_{a\in G_0} as the collection of phase spaces of the subsystems constituting the network (G, {\mathcal P}). We refer to {\mathcal P} as a phase space function. Since the state of the network should be determined completely and uniquely by the states of its subsystems, it is reasonable to take the total phase space of the network to be the product

\displaystyle{  {\mathbb{P}}(G, {\mathcal P}):= \bigsqcap_{a\in G_0} {\mathcal P}(a).  }

In the example the total phase space of the network (G,{\mathcal P}) is {\mathbb{R}}^3, while the phase space of the network (G', {\mathcal P}') is the real line {\mathbb{R}}.

Finally we need to interpret the arrows. An arrow b\xrightarrow{\gamma}a in a graph G should encode the fact that the dynamics of the subsystem associated to the node a depends on the states of the subsystem associated to the node b. To make this precise requires the notion of an ‘open system’, or ‘control system’, which we define below. It also requires a way to associate an open system to the set of arrows coming into a node/vertex a.

To a first approximation an open system (or control system, I use the two terms interchangeably) is a system of ODEs depending on parameters. I like to think of a control system geometrically: a control system on a phase space M controlled by the the phase space U is a map

F: U\times M \to TM

where TM is the tangent bundle of the space M, so that for all (u,m)\in U\times M, F(u,m) is a vector tangent to M at the point m. Given phase spaces U and M the set of all corresponding control systems forms a vector space which we denote by

\displaystyle{ \mathsf{Ctrl}(U\times M \to M)}

(More generally one can talk about the space of control systems associated with a surjective submersion Q\to M. For us, submersions of the form U\times M \to M are general enough.)

To encode the incoming arrows, we introduce the input tree I(a) (this is a very short tree, a corolla if you like). The input tree of a node a of a graph G is a directed graph whose arrows are precisely the arrows of G coming into the vertex a, but any two parallel arrows of G with target a will have disjoint sources in I(a). In the example the input tree I of the one node of a of G' is the tree

There is always a map of graphs \xi:I(a) \to G. For instance for the input tree in the example we just discussed, \xi is the map

Consequently if (G,{\mathcal P}) is a network and I(a) is an input tree of a node of G, then (I(a), {\mathcal P}\circ \xi) is also a network. This allows us to talk about the phase space {\mathbb{P}} I(a) of an input tree. In our running example,

{\mathbb{P}} I(a) = {\mathbb{R}}^2

Given a network (G,{\mathcal P}), there is a vector space \mathsf{Ctrl}({\mathbb{P}} I(a)\to {\mathbb{P}} a) of open systems associated with every node a of G.

In our running example, the vector space associated to the one node a of (G',{\mathcal P}') is

\mathsf{Ctrl}({\mathbb{R}}^2, {\mathbb{R}})  \simeq C^\infty({\mathbb{R}}^2, {\mathbb{R}})

In the same example, the network (G,{\mathcal P}) has three nodes and we associate the same vector space C^\infty({\mathbb{R}}^2, {\mathbb{R}}) to each one of them.

We then construct an interconnection map

\displaystyle{  {\mathcal{I}}: \bigsqcap_{a\in G_0} \mathsf{Ctrl}({\mathbb{P}} I(a)\to {\mathbb{P}} a) \to \Gamma (T{\mathbb{P}}(G, {\mathcal P})) }

from the product of spaces of all control systems to the space of vector fields

\Gamma (T{\mathbb{P}} (G, {\mathcal P}))

on the total phase space of the network. (We use the standard notation to denote the tangent bundle of a manifold R by TR and the space of vector fields on R by \Gamma (TR)). In our running example the interconnection map for the network (G',{\mathcal P}') is the map

\displaystyle{  {\mathcal{I}}: C^\infty({\mathbb{R}}^2, {\mathbb{R}}) \to C^\infty({\mathbb{R}}, {\mathbb{R}}), \quad f(x,u) \mapsto f(x,x).  }

The interconnection map for the network (G,{\mathcal P}) is the map

\displaystyle{  {\mathcal{I}}: C^\infty({\mathbb{R}}^2, {\mathbb{R}})^3 \to C^\infty({\mathbb{R}}^3, {\mathbb{R}}^3)}

given by

\displaystyle{  ({\mathscr{I}}(f_1,f_2, f_3))\,(x_1,x_2, x_3) = (f_1(x_1,x_2), f_2(x_2,x_1),  f_3(x_3,x_2)).  }

To summarize: a dynamical system on a network of phase spaces is the data (G, {\mathcal P}, (w_a)_{a\in G_0} ) where G=\{G_1\rightrightarrows G_0\} is a directed graph, {\mathcal P}:{\mathcal P}:G_0\to {\mathsf{PhaseSpace}} is a phase space function and (w_a)_{a\in G_0} is a collection of open systems compatible with the graph and the phase space function. The corresponding vector field on the total space of the network is obtained by interconnecting the open systems.

Dynamical systems on networks can be made into the objects of a category. Carrying this out gives us a way to associate maps of dynamical systems to combinatorial data.

The first step is to form the category of networks of phase spaces, which we call {\mathsf{Graph}}/{\mathsf{PhaseSpace}}. In this category, by definition, a morphism from a network (G,{\mathcal P}) to a network (G', {\mathcal P}') is a map of directed graphs \varphi:G\to G' which is compatible with the phase space functions:

\displaystyle{  {\mathcal P}'\circ \varphi = {\mathcal P}.  }

Using the universal properties of products it is easy to show that a map of networks \varphi: (G,{\mathcal P})\to (G',{\mathcal P}') defines a map {\mathbb{P}}\varphi of total phase spaces in the opposite direction:

\displaystyle{  {\mathbb{P}} \varphi: {\mathbb{P}} G' \to {\mathbb{P}} G.  }

In the category theory language the total phase space assignment extends to a contravariant functor

\displaystyle{ {\mathbb{P}}:  {({\mathsf{Graph}}/{\mathsf{Region}})}^{\mbox{\sf {\tiny {op}}}} \to  {\mathsf{Region}}.  }

We call this functor the total phase space functor. In our running example, the map

{\mathbb{P}}\varphi:{\mathbb{R}} = {\mathbb{P}}(G',{\mathcal P}') \to  {\mathbb{R}}^3 = {\mathbb{P}} (G,{\mathcal P})

is given by

\displaystyle{  {\mathbb{P}} \varphi (x) = (x,x,x).  }

Continuous-time dynamical systems also form a category, which we denote by \mathsf{DS}. The objects of this category are pairs consisting of a phase space and a vector field on this phase space. A morphism in this category is a smooth map of phase spaces that intertwines the two vector fields. That is:

\displaystyle{  \mathrm{Hom}_\mathsf{DS} ((M,X), (N,Y))   = \{f:M\to N \mid Df \circ X = Y\circ f\}  }

for any pair of objects (M,X), (N,Y) in \mathsf{DS}.

In general, morphisms in this category are difficult to determine explicitly. For example if (M, X) = ((a,b), \frac{d}{dt}) then a morphism from (M,X) to some dynamical system (N,Y) is simply a piece of an integral curve of the vector field Y defined on an interval (a,b). And if (M, X) = (S^1, \frac{d}{d\theta}) is the constant vector field on the circle then a morphism from (M,X) to (N,Y) is a periodic orbit of Y. Proving that a given dynamical system has a periodic orbit is usually hard.

Consequently, given a map of networks

\varphi:(G,{\mathcal P})\to (G',{\mathcal P}')

and a collection of open systems

\{w'_{a'}\}_{a'\in G'_0}

on (G',{\mathcal P}') we expect it to be very difficult if not impossible to find a collection of open systems \{w_a\}_{a\in G_0} so that

\displaystyle{  {\mathbb{P}} \varphi: ({\mathbb{P}} G', {\mathscr{I}}' (w'))\to ({\mathbb{P}} G, {\mathscr{I}} (w))  }

is a map of dynamical systems.

It is therefore somewhat surprising that there is a class of maps of graphs for which the above problem has an easy solution! The graph maps of this class are known by several different names. Following Boldi and Vigna we refer to them as graph fibrations. Note that despite what the name suggests, graph fibrations in general are not required to be surjective on nodes or edges. For example, the inclusion

is a graph fibration. We say that a map of networks

\varphi:(G,{\mathcal P})\to (G',{\mathcal P}')

is a fibration of networks if \varphi:G\to G' is a graph fibration. With some work one can show that a fibration of networks induces a pullback map

\displaystyle{  \varphi^*: \bigsqcap_{b\in G_0'} \mathsf{Ctrl}({\mathbb{P}} I(b)\to {\mathbb{P} b) \to  \bigsqcap_{a\in G_0} \mathsf{Ctrl}({\mathbb{P}}} I(a)\to {\mathbb{P}} a)  }

on the sets of tuples of the associated open systems. The main result of Dynamics on networks of manifolds is a proof that for a fibration of networks \varphi:(G,{\mathcal P})\to (G',{\mathcal P}') the maps \varphi^*, {\mathbb{P}} \varphi and the two interconnection maps {\mathcal{I}} and {\mathcal{I}}' are compatible:

Theorem. Let \varphi:(G,{\mathcal P})\to (G',{\mathcal P}') be a fibration of networks of manifolds. Then the pullback map

\displaystyle{ \varphi^*: \mathsf{Ctrl}(G',{\mathcal P}')\to \mathsf{Ctrl}(G,{\mathcal P})  }

is compatible with the interconnection maps

\displaystyle{  {\mathcal{I}}': \mathsf{Ctrl}(G',{\mathcal P}')) \to \Gamma (T{\mathbb{P}} G') }

and

\displaystyle{  {\mathcal{I}}:  (\mathsf{Ctrl}(G,{\mathcal P})) \to \Gamma (T{\mathbb{P}} G)  }

Namely for any collection w'\in \mathsf{Ctrl}(G',{\mathcal P}') of open systems on the network (G', {\mathcal P}') the diagram

commutes. In other words,

\displaystyle{ {\mathbb{P}} \varphi: ({\mathbb{P}}  (G',{\mathcal P}'), {\mathscr{I}}' (w'))\to ({\mathbb{P}} (G,  {\mathcal P}), {\mathscr{I}} (\varphi^* w')) }

is a map of continuous-time dynamical systems, a morphism in \mathsf{DS}.

At this point the pure mathematician in me is quite happy: I have a theorem! Better yet, the theorem solves the puzzle at the beginning of this post. But if you have read this far, you may well be wondering: “Ok, you told us about your theorem. Very nice. Now what?”

There is plenty to do. On the practical side of things, the continuous-time dynamical systems are much too limited for contemporary engineers. Most of the engineers I know care a lot more about hybrid systems. These kinds of systems exhibit both continuous time behavior and abrupt transitions, hence the name. For example, anti-lock breaks on a car is just that kind of a system — if a sensor detects that the car is skidding and the foot break is pressed, it starts pulsing the breaks. This is not your father’s continuous time dynamical system! Hybrid dynamical systems are very hard to understand. They have been all the rage in the engineering literature for the last 10-15 years. Sadly, most pure mathematicians never heard of them. It would be interesting to extend the theorem I am bragging about to hybrid systems.

Here is another problem that may be worth thinking about: how much of the theorem holds up to numerical simulation? My feeling is that any explicit numerical integration method will behave well. Implicit methods I am not sure about.

And finally a more general issue. John has been talking about networks quite a bit on this blog. But his networks and my networks look like very different mathematical structures. What do they have in common besides nodes and arrows? What mathematical structure are they glimpses of?


Network Theory II

12 March, 2014

 

Chemists are secretly doing applied category theory! When chemists list a bunch of chemical reactions like

C + O₂ → CO₂

they are secretly describing a ‘category’.

That shouldn’t be surprising. A category is simply a collection of things called objects together with things called morphisms going from one object to another, often written

f: x → y

The rules of a category say:

1) we can compose a morphism f: x → y and another morphism g: y → z to get an arrow gf: x → z,

2) (hg)f = h(gf), so we don’t need to bother with parentheses when composing arrows,

3) every object x has an identity morphism 1ₓ: x → x that obeys 1ₓ f = f and f 1ₓ = f.

Whenever we have a bunch of things (objects) and processes (arrows) that take one thing to another, we’re likely to have a category. In chemistry, the objects are bunches of molecules and the arrows are chemical reactions. But we can ‘add’ bunches of molecules and also add reactions, so we have something more than a mere category: we have something called a symmetric monoidal category.

My talk here, part of a series, is an explanation of this viewpoint and how we can use it to take ideas from elementary particle physics and apply them to chemistry! For more details try this free book:

• John Baez and Jacob Biamonte, A Course on Quantum Techniques for Stochastic Mechanics.

as well as this paper on the Anderson–Craciun–Kurtz theorem (discussed in my talk):

• John Baez and Brendan Fong, Quantum techniques for studying equilibrium in reaction networks.

You can also see the slides of this talk. Click on any picture in the slides, or any text in blue, and get more information!


Network Theory I

2 March, 2014

 

Here’s a video of a talk I gave last Tuesday—part of a series. You can see the slides here:

Network Theory I: electrical circuits and signal-flow graphs.

Click on items in blue, or pictures, for more information.

One reason I’m glad I gave this talk is because afterwards Jamie Vicary pointed out some very interesting consequences of the relations among signal-flow diagrams listed in my talk. It turns out they imply equations familiar from the theory of complementarity in categorical quantum mechanics!

This is the kind of mathematical surprise that makes life worthwhile for me. It seemed utterly shocking at first, but I think I’ve figured out why it happens. Now is not the time to explain… but I’ll have to do it soon, both here and in the paper that Jason Eberle are writing about control theory.

For now, besides the slides, the best place to read more about this program is here:

• Brendan Fong, A compositional approach to control theory.


Categories in Control

6 February, 2014

 

I’m visiting Erlangen from now until the end of May, since my wife got a grant to do research here. I’m trying to get a lot of papers finished. But today I’m giving a talk in the math department of the university here, which with Germanic brevity is called the Friedrich-Alexander-Universität Erlangen-Nürnberg.

You can see my slides here, or maybe even come to my talk:

Categories in control, Thursday 6 February 2014, 16:15–18:00, Mathematics Department of the FAU, in Übungsraum 1.

The title is a pun. It’s about categories in control theory, the branch of engineering that studies dynamical systems with inputs and outputs, and how to optimize their behavior.

Control theorists often describe these systems using signal-flow graphs. Here is a very rough schematic signal-flow graph, describing the all-important concept of a ‘feedback loop’:

Here is a detailed one, describing a specific device called a servo:

The device is shown on top, and the signal-flow graph describing its behavior is at bottom. For details, click on the picture.

Now, if you have a drop of category-theorist’s blood in your veins, you’ll look at this signal-flow graph and think my god, that’s a string diagram for a morphism in a monoidal category!

And you’d be right. But if you want to learn what that means, and why it matters, read my talk slides!

The slides should make sense if you’re a mathematician, but maybe not otherwise. So, here’s the executive summary. The same sort of super-abstract math that handles things like Feynman diagrams:

also handles signal-flow graphs. The details are different in important and fascinating ways, and this is what I’m mainly concerned with. But we now understand how signal-flow graphs fit into the general theory of networks. This means we can proceed to use modern math to study them—and their relation to other kinds of networks, like electrical circuit diagrams:

More talks

Thanks to the Azimuth Project team, my graduate students and many other folks, the dream of network theory as a step toward ‘green mathematics’ seems to be coming true! There’s a vast amount left to be done, so I’d have trouble convincing a skeptic, but I feel the project has turned a corner. I now feel in my bones that it’s going to work: we’ll eventually develop a language for biology and ecology based in part on category theory.

So, I think it’s a good time to explain all the various aspects of this project that have been cooking away—some quite visibly, but others on secret back burners:

• Jacob Biamonte and I have written a book on Petri nets and chemical reaction networks. You may have seen parts of this on the blog. We started this project at the Centre for Quantum Technologies, but now he’s working at the Institute for Scientific Interchange, in Turin—and collaborating with people there on various aspects of network theory.

• Brendan Fong is working with me on electrical circuits. You may know him for his posts here on chemical reaction networks. He’s now a grad student in computer science at Oxford.

• Jason Erbele, a math grad student at U.C. Riverside, is working with me on control theory. This work is the main topic of my talk—but I also sketch how it ties together with what Brendan is doing. There’s a lot more to say here.

• Tobias Fritz, a postdoc at the Perimeter Institute, is working with me on category-theoretic aspects of information theory. We published a paper on entropy with Tom Leinster, and we’ve got a followup on relative entropy that’s almost done. I should be working on it right this instant! But for now, read the series of posts here on Azimuth: Relative Entropy Part 1, Part 2 and Part 3.

• Brendan Fong has also done some great work on Bayesian networks, using ideas that connect nicely to what Tobias and I are doing.

• Tu Pham and Franciscus Rebro are working on the math that underlies all these projects: bicategories of spans.

The computer science department at Oxford is a great place for category theory and diagrammatic reasoning, thanks to the presence of Samson Abramsky, Bob Coecke and others. I’m going to visit them from February 21 to March 14. It seems like a good time to give a series of talks on this stuff. So, stay tuned! I’ll try to make slides available here.


Wormholes and Entanglement

20 January, 2014

 

An apparent contradiction in what most physicists believe about black holes—the ‘firewall problem’—is making some very good physicists reach for some very crazy-sounding ideas to find a way out. In particular, Maldacena and Susskind have come up with the idea that any pair of quantum-entangled particles is actually connected by a wormhole.

Entanglement is a spooky way for far-away particles to be correlated, but you can’t use it communicate faster than light. It’s been seen in the lab, but it’s only possible thanks to quantum mechanics. The first people to make a fuss over entanglement were Einstein, Podolsky and Rosen, back in 1935.

A wormhole is a spooky way for far-away regions of space to be connected by a kind of ‘tunnel’—but you probably can’t use it to communicate faster than light. Nobody has ever seen one, but they’re theoretically possible thanks to general relativity. The first people to make a fuss over wormholes were Einstein and Rosen, back in 1935.

So, superficially, it makes sense that there should be a connection between wormholes and entanglement. But when you learn enough physics, you’ll see that Maldacena and Susskind’s proposal sounds completely hare-brained.

But when you learn more physics—maybe more than enough?—you might decide there’s some merit to this idea after all. At the Centre for Quantum Technologies last summer, Jamie Vicary and I noticed some interesting connections between wormholes and quantum entanglement. We now have a paper out!

In it, we study quantum gravity in a universe where space is just 2-dimensional, not 3-dimensional like ours. It’s not realistic, but it has one huge advantage: there’s a working theory of what quantum gravity could be like when space is 2-dimensional, so you can calculate stuff!

So, we calculate what happens when a wormhole forms, and we show the ends look like a particle and its antiparticle (this was already known), and we note that this particle-antiparticle pair is entangled. In fact it’s completely entangled: any piece of information you might want to know about one can also be found in the other.

However, in a sense that Jamie and I make precise, this entanglement is ‘fake’. The reason is that the two ends of the wormhole are not independent things. They’re just two views of the same thing… and, technically, it doesn’t count as entanglement when something is ‘entangled with itself’. This fact is crucial to how Maldacena and Susskind want to get around the firewall problem.

For more details, try this:

Wormholes and entanglement, The n-Category Café.

This has links to other stuff, including our paper, but also some blog articles explaining the firewall problem, the paper by Maldacena and Susskind, and the original Einstein–Podolsky–Rosen and Einstein–Rosen papers (in English).

Since this quantum gravity stuff is more suited to the n-Category Café than here, I won’t enable comments here. If you want to talk, please go there. Sorry!


Quantropy (Part 4)

11 November, 2013

There’s a new paper on the arXiv:

• John Baez and Blake Pollard, Quantropy.

Blake is a physics grad student at U. C. Riverside who plans to do his thesis with me.

If you have carefully read all my previous posts on quantropy (Part 1, Part 2 and Part 3), there’s only a little new stuff here. But still, it’s better organized, and less chatty.

And in fact, Blake came up with a lot of new stuff for this paper! He studied the quantropy of the harmonic oscillator, and tweaked the analogy between statistical mechanics and quantum mechanics in an interesting way. Unfortunately, we needed to put a version of this paper on the arXiv by a deadline, and our writeup of this new work wasn’t quite ready (my fault). So, we’ll put that other stuff in a new version—or, I’m thinking now, a separate paper.

But here are two new things.

First, putting this paper on the arXiv had the usual good effect of revealing some existing work on the same topic. Joakim Munkhammar emailed me and pointed out this paper, which is free online:

• Joakim Munkhammar, Canonical relational quantum mechanics from information theory, Electronic Journal of Theoretical Physics 8 (2011), 93–108.

You’ll see it cites Garrett Lisi’s paper and pushes forward in various directions. There seems to be a typo where he writes the path integral

Z = \displaystyle{ \int e^{-\alpha S(q) } D q}

and says

In order to fit the purpose Lisi concluded that the Lagrange multiplier value \alpha \equiv 1/i \hbar. In similarity with Lisi’s approach we shall also assume that the arbitrary scaling-part of the constant \alpha is in fact 1/\hbar.

I’m pretty sure he means 1/i\hbar, given what he writes later. However, he speaks of ‘maximizing entropy’, which is not quite right for a complex-valued quantity; Blake and I prefer to give this new quantity a new name, and speak of ‘finding a stationary point of quantropy’.

But in a way these are small issues; being a mathematician, I’m more quick to spot tiny technical defects than to absorb significant new ideas, and it will take a while to really understand Munkhammar’s paper.

Second, while writing our paper, Blake and I noticed another similarity between the partition function of a classical ideal gas and the partition function of a quantum free particle. Both are given by an integral like this:

Z = \displaystyle{\int e^{-\alpha S(q) } D q }

where S is a quadratic function of q \in \mathbb{R}^n. Here n is the number of degrees of freedom for the particles in the ideal gas, or the number of time steps for a free particle on a line (where we are discretizing time). The only big difference is that

\alpha = 1/kT

for the ideal gas, but

\alpha = 1/i \hbar

for the free particle.

In both cases there’s an ambiguity in the answer! The reason is that to do this integral, we need to pick a measure D q. The obvious guess is Lebesgue measure

dq = dq_1 \cdots dq_n

on \mathbb{R}^n. But this can’t be right, on physical grounds!

The reason is that the partition function Z needs to be dimensionless, but d q has units. To correct this, we need to divide dq by some dimensionful quantity to get D q.

In the case of the ideal gas, this dimensionful quantity involves the ‘thermal de Broglie wavelength’ of the particles in the gas. And this brings Planck’s constant into the game, even though we’re not doing quantum mechanics: we’re studying the statistical mechanics of a classical ideal gas!

That’s weird and interesting. It’s not the only place where we see that classical statistical mechanics is incomplete or inconsistent, and we need to introduce some ideas from quantum physics to get sensible answers. The most famous one is the ultraviolet catastrophe. What are all rest?

In the case of the free particle, we need to divide by a quantity with dimensions of lengthn to make

dq = dq_1 \cdots dq_n

dimensionless, since each dq_i has dimensions of length. The easiest way is to introduce a length scale \Delta x and divide each dq_i by that. This is commonly done when people study the free particle. This length scale drops out of the final answer for the questions people usually care about… but not for the quantropy.

Similarly, Planck’s constant drops out of the final answer for some questions about the classical ideal gas, but not for its entropy!

So there’s an interesting question here, about what this new length scale \Delta x means, if anything. One might argue that quantropy is a bad idea, and the need for this new length scale to make it unambiguous is just proof of that. However, the mathematical analogy to quantum mechanics is so precise that I think it’s worth going a bit further out on this limb, and thinking a bit more about what’s going on.

Some weird sort of déjà vu phenomenon seems to be going on. Once upon a time, people tried to calculate the partition functions of classical systems. They discovered they were infinite or ambiguous until they introduced Planck’s constant, and eventually quantum mechanics. Then Feynman introduced the path integral approach to quantum mechanics. In this approach one is again computing partition functions, but now with a new meaning, and with complex rather than real exponentials. But these partition functions are again infinite or ambiguous… for very similar mathematical reasons! And at least in some cases, we can remove the ambiguity using the same trick as before: introducing a new constant. But then… what?

Are we stuck in an infinite loop here? What, if anything, is the meaning of this ‘second Planck’s constant’? Does this have anything to do with second quantization? (I don’t see how, but I can’t resist asking.)


The Elitzur–Vaidman Bomb-Testing Method

24 August, 2013

Quantum mechanics forces us to refine our attitude to counterfactual conditionals: questions about what would have happened if we had done something, even though we didn’t.

“What would the position of the particle be if I’d measured that… when actually I measured its momentum?” Here you’ll usually get no definite answer.

But sometimes you can use quantum mechanics to find out what would have happened if you’d done something… when classically it seems impossible!

Suppose you have a bunch of bombs. Some have a sensor that will absorb a photon you shine on it, and make the bomb explode! Others have a broken sensor, that won’t interact with the photon at all.

Can you choose some working bombs? You can tell if a bomb works by shining a photon on it. But if it works, it blows up—and then it doesn’t work anymore!

So, it sounds impossible. But with quantum mechanics you can do it. You can find some bombs that would have exploded if you had shone photons at them!

Here’s how:

Put a light that emits a single photon at A. Have the photon hit the half-silvered mirror at lower left, so it has a 50% chance of going through to the right, and a 50% chance of reflecting and going up. But in quantum mechanics, it sort of does both!

Put a bomb at B. Recombine the photon’s paths using two more mirrors. Have the two paths meet at a second half-silvered mirror at upper right. You can make it so that if the bomb doesn’t work, the photon interferes with itself and definitely goes to C, not D.

But if the bomb works, it absorbs the photon and explodes unless the photon takes the top route… in which case, when it hits the second half-silvered mirror, it has a 50% chance of going to C and a 50% chance of going to D.

So:

• If the bomb doesn’t work, the photon has a 100% chance of going to C.

• If the bomb works, there’s a 50% chance that it absorbs the photon and explodes. There’s also a 50% chance that the bomb does not explode—and then the photon is equally likely to go to either C or D. So, the photon has a 25% chance of reaching C and a 25% chance of reaching D.

So: if you see a photon at D, you know you have a working bomb… but the bomb has not exploded!

For each working bomb there’s:

• a 50% chance that it explodes,
• a 25% chance that it doesn’t explode but you can’t tell if it works,
• a 25% chance that it doesn’t explode but you can tell that it works.

This is the Elitzur–Vaidman bomb-testing method. It was invented by Avshalom Elitzur and Lev Vaidman in 1993. One year later, physicists actually did an experiment to show this idea works… but alas, not using actual bombs!

In 1996, Kwiat showed that using more clever methods, you can reduce the percentage of wasted working bombs as close to zero as you like. And pushing the idea even further, Graeme Mitchison and Richard Jozsa showed in 1999 that you can get a quantum computer to do a calculation for you without even turning it on!

This sounds amazing, but it’s really no more amazing than the bomb-testing method I’ve already described.

References

For details, read these:

• A. Elitzur and L. Vaidman, Quantum mechanical interaction-free measurements, Found. Phys. 23 (1993), 987–997.

• Paul G. Kwiat, H. Weinfurter, T. Herzog, A. Zeilinger, and M. Kasevich, Experimental realization of “interaction-free” measurements.

• Paul G. Kwiat, Interaction-free measurements.

• Graeme Mitchison and Richard Jozsa, Counterfactual computation, Proc. Roy. Soc. Lond. A457 (2001), 1175–1194.

The picture is from the Wikipedia article, which also has other references:

Elitzur–Vaidman bomb tester, Wikipedia.

Bas Spitters pointed out this category-theoretic analysis of the issue:

• Robert Furber and Bart Jacobs, Towards a categorical account of conditional probability.


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