For deterministic nested partitions with mesh tending to zero, lift the polygonal interpolation of Brownian motion by its ordinary iterated integrals. These path signatures converge almost surely to enhanced Brownian motion in every Hölder rough-path metric of exponent below $1/2$, hence in $p$-variation for $p>2$. Conditioning on the partition values gives a nested martingale for the first level and the off-diagonal second levels. Moment bounds and interpolation upgrade pointwise convergence to convergence in the rough path metric.

For $2<p<3$, the support of a measure of enhanced Brownian motion is the rough-path closure of path signatures of paths in the Cameron-Martin space of Wiener measure. Polygonal convergence gives one inclusion. For the reverse inclusion, decompose Brownian motion into its polygonal interpolation and independent pinned bridges. The bridge enhancement becomes small in probability as the mesh decreases, while every finite-dimensional neighborhood of the polygonal vertices has positive Gaussian probability. Rough path translation then gives positive probability in every ball around a prescribed control path signature.

For every $2<p<3$ and every positive radius, the canonical Brownian enhancement has positive probability of lying in the displayed rough path metric ball. This controls both displacement and area. Together with Cameron-Martin quasi-invariance of enhanced Brownian motion, it puts every lifted Cameron-Martin control in the support of enhanced Brownian motion.

Rest mass is the invariant mass $m$ in the relativistic energy-momentum relation. It equals the energy measured in the particle's rest frame divided by $c^2$.

The semidiscrete system $U^{\prime }=D_{xx}U+\alpha D_{0}U$, for real fixed $\alpha$, is contractive in the periodic discrete L2 norm. Discrete summation by parts makes the drift term's real energy contribution zero and the diffusion term nonpositive. Equivalently its Fourier amplification symbol generator has real part $-4h^{-2}{\sin }^2(\theta /2)$. This continuous-time estimate does not prove stability of an arbitrarily chosen time integrator.

For zero endpoint values and real constant drift, integration by parts gives $\frac{1}{2}d\Vert u{\Vert }_{L^2}^2/dt=-\Vert u_x{\Vert }_{L^2}^2$. The drift contributes only a zero boundary term in the real part. Thus the solution is contractive, uniformly in the drift. Multiplication by $e^{\kappa x}$ converts the equation to a heat equation with constant killing rate ${\kappa }^2$, providing existence and uniqueness as well as the direct energy estimate.

The inertial range lies between the energy-containing and viscous scales. Direct forcing and viscous dissipation are small there, and nonlinear transfer carries an approximately constant mean energy flux in a three-dimensional forward cascade. Scale similarity and universality are additional statistical hypotheses.

The dissipation range contains scales at which kinematic viscosity appreciably affects velocity gradients and converts kinetic energy into heat. In a high-Reynolds-number three-dimensional cascade its characteristic length and time are the Kolmogorov microscales. Intense vortex cores can be intermittent within this range.

The equilibrium range consists of scales below the energy-containing motions whose statistics can adjust rapidly to the local energy supply. It includes the inertial range and dissipative scales. Approximate equilibrium concerns mean transfer and viscous dissipation, not absence of intermittent spatial fluctuations.

Kolmogorov 1941 theory assumes locally universal small-scale velocity statistics set by mean viscous dissipation and kinematic viscosity at large Reynolds number. In its inertial range dimensional form, $S_p(r)\sim {ϵ}^{p/3}{r}^{p/3}$. Large-scale modulation and internal intermittency require qualifying this unconditional scaling.

The Townsend-Betchov cascade cartoon organizes inertial range vorticity into sheets by bi-axial strain, then uses sheet instability and roll-up to generate finer vortex tubes. Dissipative tube cores balance stretching with kinematic viscosity. Such differing mechanisms across scales challenge literal identical-step, memoryless cascade models without contradicting a mean energy flux.

A two-dimensional forward two-dimensional enstrophy cascade has the idealized spectrum $E(k)\propto k^{-3}$. The energy integral above a wavenumber scales as its inverse square, whereas the cumulative enstrophy integral $\int k^{2}E(k)dk$ grows logarithmically. The spectrum used here omits possible further logarithmic corrections.

The eddy turnover time estimates how long a velocity difference $\delta u_l$ takes to distort a structure of size $l$. In a local strong energy cascade, the transfer time is assumed comparable to it. A weak wave cascade can transfer energy much more slowly because a single encounter makes only a small fractional change.

At fixed outer velocity and integral scale of turbulence, the mean viscous dissipation per unit mass approaches a positive finite value as the Reynolds number becomes large. The dimensionless coefficient $C_{ϵ}=ϵ\ell /u^3$ is then of order one. This is a phenomenological property of established three-dimensional turbulence, not a consequence of taking kinematic viscosity to zero in an arbitrary smooth flow.

The quadratic nonlinear term of the Navier-Stokes equation couples three Fourier modes whose wave vectors close to a triangle. These spectral triad interactions transfer kinetic energy between wave numbers. A turbulence closure can model their triple correlations while retaining more scale information than a prescribed local power law.

For $u^{\prime }=v$, $v^{\prime }=u-u^3+\epsilon (\beta -u^2)v$, let $H=v^2/2-u^2/2+u^4/4$. On an outer periodic orbit with $H>0$, leading energy balance method requires $\beta =B(H)$, whereHere $u_{\max }^2=1+\sqrt{1+4H}$. The homoclinic orbit limit is $B(0)=4/5$. Differentiating the denominator gives a period integral diverging at $H=0$, whereas the derivative of the numerator stays finite; hence $B$ decreases immediately above zero. At large $H$, rescaling $u=H^{1/4}s$ gives $B(H)\propto \sqrt{H}$. Its first nondegenerate minimum selects a saddle-node bifurcation of periodic orbits; numerical quadrature gives ${\beta }_{\mathrm{fold}}\simeq 0.75226$. The cycle on the decreasing portion is unstable and the one on the increasing portion is stable, since the net energy drift has sign $\beta -B(H)$. Restoring an unperturbed coefficient $a^2$ in the linear restoring term multiplies the threshold by $a^2$.

If $h$ has a simple interior zero $u_0$ and $a\to 0^{+}$, the inner variable $v=h^{\prime }(u_0)(u-u_0)/a$ turns the local integral into $(1/|h^{\prime }(u_0)|){\int }_{\mathbb{R}}(1+v^2{)}^{-1}dv$. Thus a peak with vanishing width can retain finite area. Slowly varying positive regularization factors cancel between peak height and width when the numerator has the matching factor. An ordinary outer expansion can miss this contribution even though the peak is exponentially narrow.

Experimental Reconstruction of Functional Gene Transfer from the Tobacco Plastid Genome to the Nucleus by Sandra Stegemann and Ralph Bock (2006, Results, pp. 2870–2874) separated relocation of DNA from activation of a relocated gene. Engineered tobacco chloroplasts carried an nptII gene with nuclear expression signals beside an aadA gene with plastid expression signals. Selection on kanamycin recovered plants with the segment in the nuclear genome. Crosses then replaced the engineered chloroplasts with ordinary chloroplasts: the nuclear segment remained, but the plants were initially sensitive to spectinomycin and streptomycin. The transferred aadA gene was not yet functional in its new setting.

Leaf pieces were cultured on regeneration medium containing both drugs. Resistant regenerated plants supplied the functional readout. Using both drugs reduced selection of spontaneous plastid mutations conferring resistance mainly to one drug. Reciprocal crosses and selfing distinguished maternal plastid inheritance from nuclear Mendelian segregation; eight of sixteen initial candidates had the latter pattern. ^[ref]

The molecular readout came from RNA hybridization and DNA sequencing. A Northern blot detected aadA messenger RNA in activated lines, with different transcript sizes. Thermal asymmetric interlaced polymerase chain reaction recovered upstream DNA; products present in activated lines but absent in original-line and ordinary-plant controls were sequenced. The mapped rearrangements connected aadA to the upstream nuclear gene promoter. Thus resistance, inheritance, RNA detection and changed nucleotide sequences supported activation of the transferred nuclear gene, rather than resistance alone. ^[ref]

This engineered test demonstrates one route to functional nuclear gene expression after plastid DNA transfer. It does not establish a natural transfer rate, replay ancient endosymbiosis or demonstrate that the resulting protein is imported into a chloroplast. The authors explicitly identify acquisition of targeting information as an additional step not reproduced by this experiment. ^[ref]

Let $T$ be a singular endomorphism of a finite-dimensional vector space. Choose $0\ne u\in \ker T$ and a nonzero linear functional $\lambda$ vanishing on $\mathrm{im}T$, possible because that image of a linear map is proper. Then $S(x)=\lambda (x)u$ is a nonzero rank-one endomorphism with $TS=ST=0$.

A linear endomorphism is a linear map from a vector space to itself. In a finite-dimensional basis it is represented by a square matrix. It need not be invertible; an invertible linear endomorphism is an automorphism of the vector space.