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I’ve just posted to the arXiv my paper “Finite time blowup for Lagrangian modifications of the three-dimensional Euler equation“. This paper is loosely in the spirit of other recent papers of mine in which I explore how close one can get to supercritical PDE of physical interest (such as the Euler and Navier-Stokes equations), while still being able to rigorously demonstrate finite time blowup for at least some choices of initial data. Here, the PDE we are trying to get close to is the incompressible inviscid Euler equations

$\displaystyle \partial_t u + (u \cdot \nabla) u = - \nabla p$

$\displaystyle \nabla \cdot u = 0$

in three spatial dimensions, where ${u}$ is the velocity vector field and ${p}$ is the pressure field. In vorticity form, and viewing the vorticity ${\omega}$ as a ${2}$-form (rather than a vector), we can rewrite this system using the language of differential geometry as

$\displaystyle \partial_t \omega + {\mathcal L}_u \omega = 0$

$\displaystyle u = \delta \tilde \eta^{-1} \Delta^{-1} \omega$

where ${{\mathcal L}_u}$ is the Lie derivative along ${u}$, ${\delta}$ is the codifferential (the adjoint of the differential ${d}$, or equivalently the negative of the divergence operator) that sends ${k+1}$-vector fields to ${k}$-vector fields, ${\Delta}$ is the Hodge Laplacian, and ${\tilde \eta}$ is the identification of ${k}$-vector fields with ${k}$-forms induced by the Euclidean metric ${\tilde \eta}$. The equation${u = \delta \tilde \eta^{-1} \Delta^{-1} \omega}$ can be viewed as the Biot-Savart law recovering velocity from vorticity, expressed in the language of differential geometry.

One can then generalise this system by replacing the operator ${\tilde \eta^{-1} \Delta^{-1}}$ by a more general operator ${A}$ from ${2}$-forms to ${2}$-vector fields, giving rise to what I call the generalised Euler equations

$\displaystyle \partial_t \omega + {\mathcal L}_u \omega = 0$

$\displaystyle u = \delta A \omega.$

For example, the surface quasi-geostrophic (SQG) equations can be written in this form, as discussed in this previous post. One can view ${A \omega}$ (up to Hodge duality) as a vector potential for the velocity ${u}$, so it is natural to refer to ${A}$ as a vector potential operator.

The generalised Euler equations carry much of the same geometric structure as the true Euler equations. For instance, the transport equation ${\partial_t \omega + {\mathcal L}_u \omega = 0}$ is equivalent to the Kelvin circulation theorem, which in three dimensions also implies the transport of vortex streamlines and the conservation of helicity. If ${A}$ is self-adjoint and positive definite, then the famous Euler-Poincaré interpretation of the true Euler equations as geodesic flow on an infinite dimensional Riemannian manifold of volume preserving diffeomorphisms (as discussed in this previous post) extends to the generalised Euler equations (with the operator ${A}$ determining the new Riemannian metric to place on this manifold). In particular, the generalised Euler equations have a Lagrangian formulation, and so by Noether’s theorem we expect any continuous symmetry of the Lagrangian to lead to conserved quantities. Indeed, we have a conserved Hamiltonian ${\frac{1}{2} \int \langle \omega, A \omega \rangle}$, and any spatial symmetry of ${A}$ leads to a conserved impulse (e.g. translation invariance leads to a conserved momentum, and rotation invariance leads to a conserved angular momentum). If ${A}$ behaves like a pseudodifferential operator of order ${-2}$ (as is the case with the true vector potential operator ${\tilde \eta^{-1} \Delta^{-1}}$), then it turns out that one can use energy methods to recover the same sort of classical local existence theory as for the true Euler equations (up to and including the famous Beale-Kato-Majda criterion for blowup).

The true Euler equations are suspected of admitting smooth localised solutions which blow up in finite time; there is now substantial numerical evidence for this blowup, but it has not been proven rigorously. The main purpose of this paper is to show that such finite time blowup can at least be established for certain generalised Euler equations that are somewhat close to the true Euler equations. This is similar in spirit to my previous paper on finite time blowup on averaged Navier-Stokes equations, with the main new feature here being that the modified equation continues to have a Lagrangian structure and a vorticity formulation, which was not the case with the averaged Navier-Stokes equation. On the other hand, the arguments here are not able to handle the presence of viscosity (basically because they rely crucially on the Kelvin circulation theorem, which is not available in the viscous case).

In fact, three different blowup constructions are presented (for three different choices of vector potential operator ${A}$). The first is a variant of one discussed previously on this blog, in which a “neck pinch” singularity for a vortex tube is created by using a non-self-adjoint vector potential operator, in which the velocity at the neck of the vortex tube is determined by the circulation of the vorticity somewhat further away from that neck, which when combined with conservation of circulation is enough to guarantee finite time blowup. This is a relatively easy construction of finite time blowup, and has the advantage of being rather stable (any initial data flowing through a narrow tube with a large positive circulation will blow up in finite time). On the other hand, it is not so surprising in the non-self-adjoint case that finite blowup can occur, as there is no conserved energy.

The second blowup construction is based on a connection between the two-dimensional SQG equation and the three-dimensional generalised Euler equations, discussed in this previous post. Namely, any solution to the former can be lifted to a “two and a half-dimensional” solution to the latter, in which the velocity and vorticity are translation-invariant in the vertical direction (but the velocity is still allowed to contain vertical components, so the flow is not completely horizontal). The same embedding also works to lift solutions to generalised SQG equations in two dimensions to solutions to generalised Euler equations in three dimensions. Conveniently, even if the vector potential operator for the generalised SQG equation fails to be self-adjoint, one can ensure that the three-dimensional vector potential operator is self-adjoint. Using this trick, together with a two-dimensional version of the first blowup construction, one can then construct a generalised Euler equation in three dimensions with a vector potential that is both self-adjoint and positive definite, and still admits solutions that blow up in finite time, though now the blowup is now a vortex sheet creasing at on a line, rather than a vortex tube pinching at a point.

This eliminates the main defect of the first blowup construction, but introduces two others. Firstly, the blowup is less stable, as it relies crucially on the initial data being translation-invariant in the vertical direction. Secondly, the solution is not spatially localised in the vertical direction (though it can be viewed as a compactly supported solution on the manifold ${{\bf R}^2 \times {\bf R}/{\bf Z}}$, rather than ${{\bf R}^3}$). The third and final blowup construction of the paper addresses the final defect, by replacing vertical translation symmetry with axial rotation symmetry around the vertical axis (basically, replacing Cartesian coordinates with cylindrical coordinates). It turns out that there is a more complicated way to embed two-dimensional generalised SQG equations into three-dimensional generalised Euler equations in which the solutions to the latter are now axially symmetric (but are allowed to “swirl” in the sense that the velocity field can have a non-zero angular component), while still keeping the vector potential operator self-adjoint and positive definite; the blowup is now that of a vortex ring creasing on a circle.

As with the previous papers in this series, these blowup constructions do not directly imply finite time blowup for the true Euler equations, but they do at least provide a barrier to establishing global regularity for these latter equations, in that one is forced to use some property of the true Euler equations that are not shared by these generalisations. They also suggest some possible blowup mechanisms for the true Euler equations (although unfortunately these mechanisms do not seem compatible with the addition of viscosity, so they do not seem to suggest a viable Navier-Stokes blowup mechanism).

The Euler equations for three-dimensional incompressible inviscid fluid flow are

$\displaystyle \partial_t u + (u \cdot \nabla) u = - \nabla p \ \ \ \ \ (1)$

$\displaystyle \nabla \cdot u = 0$

where ${u: {\bf R} \times {\bf R}^3 \rightarrow {\bf R}^3}$ is the velocity field, and ${p: {\bf R} \times {\bf R}^3 \rightarrow {\bf R}}$ is the pressure field. For the purposes of this post, we will ignore all issues of decay or regularity of the fields in question, assuming that they are as smooth and rapidly decreasing as needed to justify all the formal calculations here; in particular, we will apply inverse operators such as ${(-\Delta)^{-1}}$ or ${|\nabla|^{-1} := (-\Delta)^{-1/2}}$ formally, assuming that these inverses are well defined on the functions they are applied to.

Meanwhile, the surface quasi-geostrophic (SQG) equation is given by

$\displaystyle \partial_t \theta + (u \cdot \nabla) \theta = 0 \ \ \ \ \ (2)$

$\displaystyle u = ( -\partial_y |\nabla|^{-1}, \partial_x |\nabla|^{-1} ) \theta \ \ \ \ \ (3)$

where ${\theta: {\bf R} \times {\bf R}^2 \rightarrow {\bf R}}$ is the active scalar, and ${u: {\bf R} \times {\bf R}^2 \rightarrow {\bf R}^2}$ is the velocity field. The SQG equations are often used as a toy model for the 3D Euler equations, as they share many of the same features (e.g. vortex stretching); see this paper of Constantin, Majda, and Tabak for more discussion (or this previous blog post).

I recently found a more direct way to connect the two equations. We first recall that the Euler equations can be placed in vorticity-stream form by focusing on the vorticity ${\omega := \nabla \times u}$. Indeed, taking the curl of (1), we obtain the vorticity equation

$\displaystyle \partial_t \omega + (u \cdot \nabla) \omega = (\omega \cdot \nabla) u \ \ \ \ \ (4)$

while the velocity ${u}$ can be recovered from the vorticity via the Biot-Savart law

$\displaystyle u = (-\Delta)^{-1} \nabla \times \omega. \ \ \ \ \ (5)$

The system (4), (5) has some features in common with the system (2), (3); in (2) it is a scalar field ${\theta}$ that is being transported by a divergence-free vector field ${u}$, which is a linear function of the scalar field as per (3), whereas in (4) it is a vector field ${\omega}$ that is being transported (in the Lie derivative sense) by a divergence-free vector field ${u}$, which is a linear function of the vector field as per (5). However, the system (4), (5) is in three dimensions whilst (2), (3) is in two spatial dimensions, the dynamical field is a scalar field ${\theta}$ for SQG and a vector field ${\omega}$ for Euler, and the relationship between the velocity field and the dynamical field is given by a zeroth order Fourier multiplier in (3) and a ${-1^{th}}$ order operator in (5).

However, we can make the two equations more closely resemble each other as follows. We first consider the generalisation

$\displaystyle \partial_t \omega + (u \cdot \nabla) \omega = (\omega \cdot \nabla) u \ \ \ \ \ (6)$

$\displaystyle u = T (-\Delta)^{-1} \nabla \times \omega \ \ \ \ \ (7)$

where ${T}$ is an invertible, self-adjoint, positive-definite zeroth order Fourier multiplier that maps divergence-free vector fields to divergence-free vector fields. The Euler equations then correspond to the case when ${T}$ is the identity operator. As discussed in this previous blog post (which used ${A}$ to denote the inverse of the operator denoted here as ${T}$), this generalised Euler system has many of the same features as the original Euler equation, such as a conserved Hamiltonian

$\displaystyle \frac{1}{2} \int_{{\bf R}^3} u \cdot T^{-1} u,$

the Kelvin circulation theorem, and conservation of helicity

$\displaystyle \int_{{\bf R}^3} \omega \cdot T^{-1} u.$

Also, if we require ${\omega}$ to be divergence-free at time zero, it remains divergence-free at all later times.

Let us consider “two-and-a-half-dimensional” solutions to the system (6), (7), in which ${u,\omega}$ do not depend on the vertical coordinate ${z}$, thus

$\displaystyle \omega(t,x,y,z) = \omega(t,x,y)$

and

$\displaystyle u(t,x,y,z) = u(t,x,y)$

but we allow the vertical components ${u_z, \omega_z}$ to be non-zero. For this to be consistent, we also require ${T}$ to commute with translations in the ${z}$ direction. As all derivatives in the ${z}$ direction now vanish, we can simplify (6) to

$\displaystyle D_t \omega = (\omega_x \partial_x + \omega_y \partial_y) u \ \ \ \ \ (8)$

where ${D_t}$ is the two-dimensional material derivative

$\displaystyle D_t := \partial_t + u_x \partial_x + u_y \partial_y.$

Also, divergence-free nature of ${\omega,u}$ then becomes

$\displaystyle \partial_x \omega_x + \partial_y \omega_y = 0$

and

$\displaystyle \partial_x u_x + \partial_y u_y = 0. \ \ \ \ \ (9)$

In particular, we may (formally, at least) write

$\displaystyle (\omega_x, \omega_y) = (\partial_y \theta, -\partial_x \theta)$

for some scalar field ${\theta(t,x,y,z) = \theta(t,x,y)}$, so that (7) becomes

$\displaystyle u = T ( (- \Delta)^{-1} \partial_y \omega_z, - (-\Delta^{-1}) \partial_x \omega_z, \theta ). \ \ \ \ \ (10)$

The first two components of (8) become

$\displaystyle D_t \partial_y \theta = \partial_y \theta \partial_x u_x - \partial_x \theta \partial_y u_x$

$\displaystyle - D_t \partial_x \theta = \partial_y \theta \partial_x u_y - \partial_x \theta \partial_y u_y$

which rearranges using (9) to

$\displaystyle \partial_y D_t \theta = \partial_x D_t \theta = 0.$

Formally, we may integrate this system to obtain the transport equation

$\displaystyle D_t \theta = 0, \ \ \ \ \ (11)$

Finally, the last component of (8) is

$\displaystyle D_t \omega_z = \partial_y \theta \partial_x u_z - \partial_x \theta \partial_y u_z. \ \ \ \ \ (12)$

At this point, we make the following choice for ${T}$:

$\displaystyle T ( U_x, U_y, \theta ) = \alpha (U_x, U_y, \theta) + (-\partial_y |\nabla|^{-1} \theta, \partial_x |\nabla|^{-1} \theta, 0) \ \ \ \ \ (13)$

$\displaystyle + P( 0, 0, |\nabla|^{-1} (\partial_y U_x - \partial_x U_y) )$

where ${\alpha > 0}$ is a real constant and ${Pu := (-\Delta)^{-1} (\nabla \times (\nabla \times u))}$ is the Leray projection onto divergence-free vector fields. One can verify that for large enough ${\alpha}$, ${T}$ is a self-adjoint positive definite zeroth order Fourier multiplier from divergence free vector fields to divergence-free vector fields. With this choice, we see from (10) that

$\displaystyle u_z = \alpha \theta - |\nabla|^{-1} \omega_z$

so that (12) simplifies to

$\displaystyle D_t \omega_z = - \partial_y \theta \partial_x |\nabla|^{-1} \omega_z + \partial_x \theta \partial_y |\nabla|^{-1} \omega_z.$

This implies (formally at least) that if ${\omega_z}$ vanishes at time zero, then it vanishes for all time. Setting ${\omega_z=0}$, we then have from (10) that

$\displaystyle (u_x,u_y,u_z) = (-\partial_y |\nabla|^{-1} \theta, \partial_x |\nabla|^{-1} \theta, \alpha \theta )$

and from (11) we then recover the SQG system (2), (3). To put it another way, if ${\theta(t,x,y)}$ and ${u(t,x,y)}$ solve the SQG system, then by setting

$\displaystyle \omega(t,x,y,z) := ( \partial_y \theta(t,x,y), -\partial_x \theta(t,x,y), 0 )$

$\displaystyle \tilde u(t,x,y,z) := ( u_x(t,x,y), u_y(t,x,y), \alpha \theta(t,x,y) )$

then ${\omega,\tilde u}$ solve the modified Euler system (6), (7) with ${T}$ given by (13).

We have ${T^{-1} \tilde u = (0, 0, \theta)}$, so the Hamiltonian ${\frac{1}{2} \int_{{\bf R}^3} \tilde u \cdot T^{-1} \tilde u}$ for the modified Euler system in this case is formally a scalar multiple of the conserved quantity ${\int_{{\bf R}^2} \theta^2}$. The momentum ${\int_{{\bf R}^3} x \cdot \tilde u}$ for the modified Euler system is formally a scalar multiple of the conserved quantity ${\int_{{\bf R}^2} \theta}$, while the vortex stream lines that are preserved by the modified Euler flow become the level sets of the active scalar that are preserved by the SQG flow. On the other hand, the helicity ${\int_{{\bf R}^3} \omega \cdot T^{-1} \tilde u}$ vanishes, and other conserved quantities for SQG (such as the Hamiltonian ${\int_{{\bf R}^2} \theta |\nabla|^{-1} \theta}$) do not seem to correspond to conserved quantities of the modified Euler system. This is not terribly surprising; a low-dimensional flow may well have a richer family of conservation laws than the higher-dimensional system that it is embedded in.

As in the previous post, all computations here are at the formal level only.

In the previous blog post, the Euler equations for inviscid incompressible fluid flow were interpreted in a Lagrangian fashion, and then Noether’s theorem invoked to derive the known conservation laws for these equations. In a bit more detail: starting with Lagrangian space ${{\cal L} = ({\bf R}^n, \hbox{vol})}$ and Eulerian space ${{\cal E} = ({\bf R}^n, \eta, \hbox{vol})}$, we let ${M}$ be the space of volume-preserving, orientation-preserving maps ${\Phi: {\cal L} \rightarrow {\cal E}}$ from Lagrangian space to Eulerian space. Given a curve ${\Phi: {\bf R} \rightarrow M}$, we can define the Lagrangian velocity field ${\dot \Phi: {\bf R} \times {\cal L} \rightarrow T{\cal E}}$ as the time derivative of ${\Phi}$, and the Eulerian velocity field ${u := \dot \Phi \circ \Phi^{-1}: {\bf R} \times {\cal E} \rightarrow T{\cal E}}$. The volume-preserving nature of ${\Phi}$ ensures that ${u}$ is a divergence-free vector field:

$\displaystyle \nabla \cdot u = 0. \ \ \ \ \ (1)$

If we formally define the functional

$\displaystyle J[\Phi] := \frac{1}{2} \int_{\bf R} \int_{{\cal E}} |u(t,x)|^2\ dx dt = \frac{1}{2} \int_R \int_{{\cal L}} |\dot \Phi(t,x)|^2\ dx dt$

then one can show that the critical points of this functional (with appropriate boundary conditions) obey the Euler equations

$\displaystyle [\partial_t + u \cdot \nabla] u = - \nabla p$

$\displaystyle \nabla \cdot u = 0$

for some pressure field ${p: {\bf R} \times {\cal E} \rightarrow {\bf R}}$. As discussed in the previous post, the time translation symmetry of this functional yields conservation of the Hamiltonian

$\displaystyle \frac{1}{2} \int_{{\cal E}} |u(t,x)|^2\ dx = \frac{1}{2} \int_{{\cal L}} |\dot \Phi(t,x)|^2\ dx;$

the rigid motion symmetries of Eulerian space give conservation of the total momentum

$\displaystyle \int_{{\cal E}} u(t,x)\ dx$

and total angular momentum

$\displaystyle \int_{{\cal E}} x \wedge u(t,x)\ dx;$

and the diffeomorphism symmetries of Lagrangian space give conservation of circulation

$\displaystyle \int_{\Phi(\gamma)} u^*$

for any closed loop ${\gamma}$ in ${{\cal L}}$, or equivalently pointwise conservation of the Lagrangian vorticity ${\Phi^* \omega = \Phi^* du^*}$, where ${u^*}$ is the ${1}$-form associated with the vector field ${u}$ using the Euclidean metric ${\eta}$ on ${{\cal E}}$, with ${\Phi^*}$ denoting pullback by ${\Phi}$.

It turns out that one can generalise the above calculations. Given any self-adjoint operator ${A}$ on divergence-free vector fields ${u: {\cal E} \rightarrow {\bf R}}$, we can define the functional

$\displaystyle J_A[\Phi] := \frac{1}{2} \int_{\bf R} \int_{{\cal E}} u(t,x) \cdot A u(t,x)\ dx dt;$

as we shall see below the fold, critical points of this functional (with appropriate boundary conditions) obey the generalised Euler equations

$\displaystyle [\partial_t + u \cdot \nabla] Au + (\nabla u) \cdot Au= - \nabla \tilde p \ \ \ \ \ (2)$

$\displaystyle \nabla \cdot u = 0$

for some pressure field ${\tilde p: {\bf R} \times {\cal E} \rightarrow {\bf R}}$, where ${(\nabla u) \cdot Au}$ in coordinates is ${\partial_i u_j Au_j}$ with the usual summation conventions. (When ${A=1}$, ${(\nabla u) \cdot Au = \nabla(\frac{1}{2} |u|^2)}$, and this term can be absorbed into the pressure ${\tilde p}$, and we recover the usual Euler equations.) Time translation symmetry then gives conservation of the Hamiltonian

$\displaystyle \frac{1}{2} \int_{{\cal E}} u(t,x) \cdot A u(t,x)\ dx.$

If the operator ${A}$ commutes with rigid motions on ${{\cal E}}$, then we have conservation of total momentum

$\displaystyle \int_{{\cal E}} Au(t,x)\ dx$

and total angular momentum

$\displaystyle \int_{{\cal E}} x \wedge Au(t,x)\ dx,$

and the diffeomorphism symmetries of Lagrangian space give conservation of circulation

$\displaystyle \int_{\Phi(\gamma)} (Au)^*$

or pointwise conservation of the Lagrangian vorticity ${\Phi^* \theta := \Phi^* d(Au)^*}$. These applications of Noether’s theorem proceed exactly as the previous post; we leave the details to the interested reader.

One particular special case of interest arises in two dimensions ${n=2}$, when ${A}$ is the inverse derivative ${A = |\nabla|^{-1} = (-\Delta)^{-1/2}}$. The vorticity ${\theta = d(Au)^*}$ is a ${2}$-form, which in the two-dimensional setting may be identified with a scalar. In coordinates, if we write ${u = (u_1,u_2)}$, then

$\displaystyle \theta = \partial_{x_1} |\nabla|^{-1} u_2 - \partial_{x_2} |\nabla|^{-1} u_1.$

Since ${u}$ is also divergence-free, we may therefore write

$\displaystyle u = (- \partial_{x_2} \psi, \partial_{x_1} \psi )$

where the stream function ${\psi}$ is given by the formula

$\displaystyle \psi = |\nabla|^{-1} \theta.$

If we take the curl of the generalised Euler equation (2), we obtain (after some computation) the surface quasi-geostrophic equation

$\displaystyle [\partial_t + u \cdot \nabla] \theta = 0 \ \ \ \ \ (3)$

$\displaystyle u = (-\partial_{x_2} |\nabla|^{-1} \theta, \partial_{x_1} |\nabla|^{-1} \theta).$

This equation has strong analogies with the three-dimensional incompressible Euler equations, and can be viewed as a simplified model for that system; see this paper of Constantin, Majda, and Tabak for details.

Now we can specialise the general conservation laws derived previously to this setting. The conserved Hamiltonian is

$\displaystyle \frac{1}{2} \int_{{\bf R}^2} u\cdot |\nabla|^{-1} u\ dx = \frac{1}{2} \int_{{\bf R}^2} \theta \psi\ dx = \frac{1}{2} \int_{{\bf R}^2} \theta |\nabla|^{-1} \theta\ dx$

(a law previously observed for this equation in the abovementioned paper of Constantin, Majda, and Tabak). As ${A}$ commutes with rigid motions, we also have (formally, at least) conservation of momentum

$\displaystyle \int_{{\bf R}^2} Au\ dx$

(which up to trivial transformations is also expressible in impulse form as ${\int_{{\bf R}^2} \theta x\ dx}$, after integration by parts), and conservation of angular momentum

$\displaystyle \int_{{\bf R}^2} x \wedge Au\ dx$

(which up to trivial transformations is ${\int_{{\bf R}^2} \theta |x|^2\ dx}$). Finally, diffeomorphism invariance gives pointwise conservation of Lagrangian vorticity ${\Phi^* \theta}$, thus ${\theta}$ is transported by the flow (which is also evident from (3). In particular, all integrals of the form ${\int F(\theta)\ dx}$ for a fixed function ${F}$ are conserved by the flow.

The Euler equations for incompressible inviscid fluids may be written as

$\displaystyle \partial_t u + (u \cdot \nabla) u = -\nabla p$

$\displaystyle \nabla \cdot u = 0$

where ${u: [0,T] \times {\bf R}^n \rightarrow {\bf R}^n}$ is the velocity field, and ${p: [0,T] \times {\bf R}^n \rightarrow {\bf R}}$ is the pressure field. To avoid technicalities we will assume that both fields are smooth, and that ${u}$ is bounded. We will take the dimension ${n}$ to be at least two, with the three-dimensional case ${n=3}$ being of course especially interesting.

The Euler equations are the inviscid limit of the Navier-Stokes equations; as discussed in my previous post, one potential route to establishing finite time blowup for the latter equations when ${n=3}$ is to be able to construct “computers” solving the Euler equations, which generate smaller replicas of themselves in a noise-tolerant manner (as the viscosity term in the Navier-Stokes equation is to be viewed as perturbative noise).

Perhaps the most prominent obstacles to this route are the conservation laws for the Euler equations, which limit the types of final states that a putative computer could reach from a given initial state. Most famously, we have the conservation of energy

$\displaystyle \int_{{\bf R}^n} |u|^2\ dx \ \ \ \ \ (1)$

(assuming sufficient decay of the velocity field at infinity); thus for instance it would not be possible for a computer to generate a replica of itself which had greater total energy than the initial computer. This by itself is not a fatal obstruction (in this paper of mine, I constructed such a “computer” for an averaged Euler equation that still obeyed energy conservation). However, there are other conservation laws also, for instance in three dimensions one also has conservation of helicity

$\displaystyle \int_{{\bf R}^3} u \cdot (\nabla \times u)\ dx \ \ \ \ \ (2)$

and (formally, at least) one has conservation of momentum

$\displaystyle \int_{{\bf R}^3} u\ dx$

and angular momentum

$\displaystyle \int_{{\bf R}^3} x \times u\ dx$

(although, as we shall discuss below, due to the slow decay of ${u}$ at infinity, these integrals have to either be interpreted in a principal value sense, or else replaced with their vorticity-based formulations, namely impulse and moment of impulse). Total vorticity

$\displaystyle \int_{{\bf R}^3} \nabla \times u\ dx$

is also conserved, although it turns out in three dimensions that this quantity vanishes when one assumes sufficient decay at infinity. Then there are the pointwise conservation laws: the vorticity and the volume form are both transported by the fluid flow, while the velocity field (when viewed as a covector) is transported up to a gradient; among other things, this gives the transport of vortex lines as well as Kelvin’s circulation theorem, and can also be used to deduce the helicity conservation law mentioned above. In my opinion, none of these laws actually prohibits a self-replicating computer from existing within the laws of ideal fluid flow, but they do significantly complicate the task of actually designing such a computer, or of the basic “gates” that such a computer would consist of.

Below the fold I would like to record and derive all the conservation laws mentioned above, which to my knowledge essentially form the complete set of known conserved quantities for the Euler equations. The material here (although not the notation) is drawn from this text of Majda and Bertozzi.